Measurement-only topological quantum computation

ABSTRACT

Measurement-only topological quantum computation using both projective and interferometrical measurement of topological charge is described. Various issues that would arise when realizing it in fractional quantum Hall systems are discussed.

CROSS-REFERENCE TO RELATED APPLICATIONS

This application is a continuation of U.S. patent application Ser. No. 13/465,374, filed May 7, 2012, which is a continuation of U.S. patent application Ser. No. 12/187,850, filed Aug. 7, 2008, which claims the benefit of U.S. Provisional Patent Application No. 60/970,652, filed Sep. 7, 2007. The entire disclosures of these applications are incorporation herein by referenced.

BACKGROUND

Since von Neumann's axiomatization of quantum mechanics in the 1930s, measurement has been a kind of stepchild to unitary evolution. s the link between quantum mechanics and the classical world, measurement has attracted considerable skeptical scrutiny from philosophers. In the domain of quantum computation, also, measurement is often dreaded: an act which may project out degrees of computational freedom and potentially decoheres useful quantum correlations. However, from a pragmatic condensed matter perspective, measurement is a co-equal pillar of quantum mechanics and a tool to be exploited. In mathematics, also, unitaries and projectors are dual. Bott periodicity states that loops of unitaries correspond to a projector and a loop of projectors corresponds to a unitary. This perspective links the Chern class in the bulk to the energy current at the edge in the integral quantum Hall systems.

It is a familiar idea that measurement can stop something from happening, e.g., the “quantum watched pot effect” of Zeno's paradox. It is less familiar that measurement can elicit an intended evolution of states. To get an idea for how this might work, consider adiabatic evolution of a vector Ψ in a degenerate ground state manifold of a Hamiltonian H. Perturbing the Hamiltonian in time, H(t), while leaving the degeneracy k intact, will evolve Ψ in time according to the canonical connection on the “tautological bundle” over the Grassmann of k-planes. A discrete description of this evolution amounts to moving the k-plane slightly, leaving Ψ behind, and then projecting Ψ orthogonally back into the moved plane and repeating. It is well known that adiabatic evolution can affect the general unitary on the ground space k-plane, so a composition of projections (“measurements”) suffices to simulate unitary evolution in this simple example. The example amounts to the “quantum watched pot effect” where the “pot” is not holding still but evolving.

The preceding example suggests that a quantum state can be deliberately nudged along by a sequence of measurements as an alternative to (directly) constructing a unitary evolution of a state's underlying degrees of freedom. In quantum computation, the accuracy of the evolution is paramount. A principle advantage of the topological model is that the unitaries corresponding to braid representations are essentially exact (topologically protected). However, the topological model also anticipates measuring states in the basis of “topological charge,” and this basis is also topologically protected. Furthermore, the prime tool for such measurements, quasiparticle interferometry, is rapidly developing in both theory and experiment. An important result is that the operation of Fabry-Perot interferometers in the Fractional Quantum Hall (FQH) context produce density matrices which converge exponentially fast to projection onto charge sectors (plus an additional projection “severing charge lines” running from the interior to exterior of the interferometer). With this tool in hand, it makes sense to ask whether we can organize universal quantum computation, again in the context of FQH fluids, as a sequence of interferometrical measurements, rather than as an exercise in braiding of quasiparticles, which has been the hypothetical paradigm since 2000. Actually, the older paradigm also required a bit of measurement, presumably interferometry, to properly initialize the system and then to measure its computational output. What we find is the topological model of computation can, in fact, be disencumbered from the necessity of braiding. From a technological point of view, this may be an important advance since much has been done experimentally with interferometers.

At some level, it should not be a surprise that interferometry can substitute for braiding. After all, an interferometer does braid a stream of quasiparticles running along the system edge (and tunneling across junctions) around other quasiparticles fixed in the bulk. At bottom, the procedures look similar. However, a little more thought turns up a conundrum: How can an operation which reduces rank (projection) simulate an operation which has full rank (unitary evolution)? The crucial answer is that we do not attempt to simulate the braid evolution of the general state vector Ψ through measurements. We only need to simulate the effect of the braid evolution on our initial state Ψ₀. That is all we ever ask of a quantum computer: to faithfully evolve a known initial state.

SUMMARY

We call the approach to quantum computation based on the measurement of topological charge “measurement-only topological quantum computation” (MOTQC). There are two flavors of MOTQC. The first is to assume that there exists an observable which can be applied to the topological medium which executes precisely von Neumann orthogonal measurements projecting into topological charge sectors. The second, takes a perhaps more pragmatic view and assumes that topological charge will be measured by quasiparticle interferometers. Mathematically, this is known to amount, up to exponentially small corrections, to von Neumann projection to charge sectors together with “severing of charge lines” connecting the inside to the outside of the interferometer. The details of the MOTQC protocols, of course, depend on the topological medium doing the computation. Our background models are certain FQH states and the non-Abelian Chern-Simons field theories which are believed to underlie them. One key example is the ν=5/2 FQH state, which experimentally is found to be the most robust FQH state beyond the lowest Landau level. Here it is expected that the non-Abelian features are captured by the Ising CFT. In this special case we argue that the two types of topological charge measurement agree, because severing charge lines is immaterial, so one simply uses the simpler projective measurement version of MOTQC. In a second prominent example, where the CFT is the

₃-Parafermions, proposed to describe the ν=12/5 plateau, we will see that there are substantial changes to the protocols passing between the two flavors. Our analysis is quite general, but these two examples will be of the greatest current interest.

Anyonic state “teleportation” may be achieved using “forced measurement” and it is applied to generate braiding transformations from an adaptive series of (non-demolitional) measurements. Forced measurement is an important primitive for us. It is an adaptive sequence of measurements (of either flavor, depending on our context) which probabilistically assures an outcome equivalent to the one “desired” outcome of the initial measurement of the sequence. Herein, we will review certain results in the density matrix formalism necessary for extension to the “second flavor,” and describe how to do all the steps when using interferometry measurements of topological charge. The reader who is not familiar with “teleportation” from the computer science side of this subject should not be concerned; it amounts merely to moving an anyon along a world line which is not monotone in time.

Topological protection of the qubits is provided naturally at the hardware level by the non-local state space in non-Abelian FQH systems. But error protection is not enough; one must be able to perform computational gate. This was originally conceived to be implemented through physical braiding of anyons in the adiabatic regime. One thing that is new here is that the effect of braiding on the protected Hilbert space, the internal degrees of freedom of the anyonic system, will be “simulated” on the initial state Ψ₀ by an adaptive sequence of non-demolitional measurements. Conceptually, the interferometric measurements will be organized into “forced measurement” teleportations, and then braids; however, the key physical primitive is interferometry. It will be seen that even a simple computation will involve patterning and re-patterning the edge of the FQH fluid (or medium) to form and reform various interferometers. The interferometers cannot be built once and for all since they must necessarily occupy overlapping regions of the fluid.

Topological quantum computation (TQC) does not rely on physically braiding the computational anyons. Quasiparticle interferometry is now accomplished in several labs, most recently by Willet, et. al., in the ν=5/2 regime.

Our protocols require that the collective topological charge be measured for a maximum of 8 neighboring anyons. In the ν=5/2 (Ising) context, our protocols can be combined with earlier strategies to extend MOTQC to achieve universality. In the ν=12/5 (

₃-Parafermions) context, universality is immediate from the reduction of braiding to interferometry.

Clearly, since interferometry is a projection together with decoherence (“severing charge lines”), the process studied here is more complicated than for projective measurement. The decoherence inherent in the process has the effect of projection to the trivial sector of the overall topological charge of the target system anyons exterior to the interferometry region, which sounds quite risky for the quantum information they contain. Fortunately, when examined in detail, this additional projection does not harm the encoded computational state information, so the situation can be salvaged.

BRIEF DESCRIPTION OF THE DRAWINGS

FIGS. 1A-1B depict topological aspects of an operator determined by the spatial configuration of a physical process encoded in their diagrammatic representation. Two topologically inequivalent spatial configurations for measuring the collective topological charge of anyons 1 and 4 in the presence of anyons 2 and 3 are shown here, with FIG. 1A corresponding to Eq. (2.8), and FIG. 1B corresponding to Eq. (2.9). The ovals delineate areas inside which the collective topological charge of anyons is being measured.

FIGS. 2A-2C depict projective topological charge measurements of pairs of anyons 1 and 2 (FIG. 2A), 2 and 3 (FIG. 2B), and 2 and 4 (FIG. 2C), used to implement forced projective measurement anyonic state teleportation, which is used to produce braiding transformations as in Eqs. (4.7-4.16). The ovals delineate the areas for which the contained collective topological charge is being measured.

FIGS. 3A-3F depict interferometry topological charge measurements of pairs of anyons 1 and 2 (FIG. 3A); 2 and 3 (FIG. 3B); and 2 and 4 (FIG. 3C), used together with measurements of multiple anyons 1, 3, . . . (FIG. 3D); 1, 4, . . . (FIG. 3E); and 3, 4, . . . (FIG. 3F); where “ . . . ” represent the additional anyons (not shown) that also comprise the state ρ₀, to implement forced interferometry measurement anyonic state teleportation, which is used to produce braiding transformations as in Eqs. (4.18-4.27). The ovals represent the interferometry loop of the probe anyons, delineating the areas for which the contained collective topological charge is being measured. The grey bars represent a “safe” choice of entry/exit paths of probe anyons.

FIG. 4 depicts a quasi-linear array of stationary anyons used for measurement-only topological quantum computing. Entanglement resource pairs of anyons (denoted by X's) are situated between adjacent pairs of computational anyons (denoted by dots) to facilitate measurement generated braiding transformations used to implement computational gates on the topological qubits.

FIG. 5 depicts measurement-only topological quantum computation in fractional quantum Hall systems, where a quasi-linear array of stationary anyons is set up, and double point-contact interferometers are used to perform interferometrical topological charge measurements. Here we show a section of an array in a Hall bar with three interferometers measuring the topological charges of different pairs of anyons. The bulk of the FQH fluid is the grey area, and its edge chirality is indicated by the arrows. The computational anyons are denoted here by dots, and the entanglement resource anyons are denoted by Xs. The FQH edge is deformed into the bulk by depleting the Hall fluid (white areas) in order to construct interferometers enveloping anyons to be measured. Once a measurement is completed, the edge protrusions are retracted, destructing the no longer needed interferometer. Tunneling across the point contacts is indicated by dashed lines. The “arms” reaching in from the side carry the incoming and outgoing current of probe anyons along their edges, and thus represent the entry/exit paths of probe anyons depicted abstractly in FIG. 3. Topological charge measurement outcomes are distinguished by the observed values of current that tunnels an interferometer. For one of the interferometers, we show where leads are attached (black squares) to a Hall bar to measure the tunneling current across the interferometer.

FIG. 6 is a block diagram showing an exemplary computing environment in which aspects of the invention may be implemented.

DETAILED DESCRIPTION OF ILLUSTRATIVE EMBODIMENTS Topological Charge Measurement

Since topological charge measurement is the foundation upon which the measurement-only approach to TQC is based, we begin by describing it in detail. We use a diagrammatic representation of anyonic states and operators acting on them, as described by an anyon model. These diagrams encode the purely topological properties of anyons, independent of any particular physical representation.

Projective Measurement

Given a complete set of Hermitian, orthogonal projection operators {π_(c)} corresponding to the eigenstates of some observable, the projective measurement of this observable for a state |Ψ

will have outcome c with probability

Prob(c)=

Ψ|πc|Ψ

  (2.1)

and project (with re-normalization) the state into the corresponding subspace

$\begin{matrix} {\left. {\left| \Psi \right.\rangle}\mapsto{\Pi_{c}\left\lbrack \left| \Psi \right.\rangle \right.} \right\rbrack = \frac{\left. \Pi_{c} \middle| \Psi \right.\rangle}{\sqrt{\langle{\Psi {\Pi_{c}}}\rangle}}} & (2.2) \end{matrix}$

In an abuse of notation, we use π_(c) to denote both the map to the post-measurement state for outcome c and the projector acting on the Hilbert space, but the meaning should be clear from context. Using the density matrix formalism to describe states, the corresponding probability and projection of the state ρ for measurement outcome c are

$\begin{matrix} {{{Prob}(c)} = {\overset{\sim}{Tr}\left\lbrack {\Pi_{c}\rho} \right\rbrack}} & (2.3) \\ {\left. \rho\mapsto{\Pi_{c}\lbrack\rho\rbrack} \right. = {\frac{\Pi_{c}{\rho\Pi}_{c}}{\overset{\sim}{Tr}\left\lbrack {\Pi_{c}\rho} \right\rbrack}.}} & (2.4) \end{matrix}$

Within the diagrammatic formalism, the projector onto definite internal vertex state μ and collective charge c of two anyons 1 and 2 (numbered from left to right) carrying definite charges a and b respectively is

For measurements that can distinguish the collective charge c, but not the different internal vertex states, the projector is given by

In the rest of this paper, we will only consider anyon models with no fusion multiplicities (i.e., the fusion coefficients are either N_(ab) ^(c)=0 or 1, depending on whether the fusion is allow), so we can drop the internal vertex label. This subset encompasses all of the most physically relevant anyon models, and generalizing the methods and results to apply when there are fusion multiplicities is straightforward, so this is not a severe restriction.

The projector onto collective topological charge c of n anyons is

When one writes operators, such as these topological charge projectors, in anionic systems, one must also specify their topological configuration with respect to other anyons in the system, which is determined by the physical process they represent. This configuration information is explicitly contained in the diagrammatic representation of operators, but left explicit in the symbolic shorthand (e.g., π_(c)) used to represent it. For example, when considering four anyons, we write

to represent the projection corresponding to a charge measurement with the spatial configuration in FIG. 1(a), whereas corresponds to the spatial configuration in FIG. 1(b). Clearly the two corresponding spatial configurations can be inferred from the diagrammatic representation of operators. When the topological configuration is left implicit, we typically the obvious simplest configuration (i.e. the planar diagrams arising from paths), but when its important we will explicitly describe the intended spatial configuration.

Specifying a region inside which the collective topological charge is measured, in FIG. 1, is like specifying a path through which the measured anyons are together to fuse into a single definite collective charge, and then separated returned back to their original positions. In fact, given the ability to somehow anyons into their original configuration of localized anyonic charges after them, this would be one possible way to perform such a charge measurement. However, there are other ways to measure the collective charge of anyons without fusing them, or even moving them at all, such as interferometrical measurement. Hence, the spatial configuration of a measurement gives the projection a path-like quality, even if the measured anyons are not actually moved. emphasize that this means the diagrams in these projection operators should interpreted as the worldlines of the anyons. Rather, these diagrams only represents the fact that the anyonic state is somehow being projected into a subspace which has a definite combined topological charge. If one were to fuse anyons together, but not split them up again, this would produce a single anyon of definite topological charge c with Prob (c) as given in Eqs. (2.1, 2.3), but the projection is instead into a state space with fewer anyons.

The physical measurement in topological systems that give rise to projective topological charge measurements of this sort potentially include: Wilson loop measurements (enclosing the measured quasiparticles) in lattice models, energy splitting measurements in fractional quantum Hall (FQH) and possibly other systems, and (the asymptotic limit of) interferometry measurements when the collective charge measurement outcome c is an Abelian charge. While motion (or something related to it) of quasiparticles may still occur in such measurements, it is something manifestly different from the braiding of the computational anyons. In particular, while Wilson loop measurements can be related to moving quasiparticles, does not move the quasiparticles enclosed inside the loop; the motion for energy splitting measurements simply brings the quasiparticles into closer proximity and then returns them to their original location (without braiding of any sort); and for interferometry, there is of course a beam of moving probe quasiparticles interfering around the measured quasiparticles, but the measured quasiparticles are not moved at all.

When performing topological charge measurements, one must be careful to avoid carrying them out in a manner or configuration that results in undesired effects on the anyonic charge correlations of the system, such as the introduction of unintentional charge entanglement or decoherence of charge entanglement that encodes relevant information.

Interferometry Measurement

In contrast with projective measurement, interferometrical measurement of topological charge is not quite as simple and requires a density matrix formulation. We will assume the anyonic interferometer is of one of two known forms, i.e., either an idealized Mach-Zehnder or Fabry-Pérot type. The Fabry-Pérot type interferometer has been experimentally realized in FQH systems using the double-point contact geometry in the weak tunneling regime. The construction of Mach-Zehnder interferometers in FQH systems unavoidably requires one of the detectors and drains to be situated inside the central interferometry region. This results in the accumulation of probe anyons which changes the topological charge contained inside this region, rendering it incapable of measuring a target charge, and hence useless for topological quantum computation (in any form).

To describe the effects of interferometry, we split the anyons of the system into three categories: the target anyons 1, . . . , n, located inside the interferometry region (collectively written as A); the (non-probe) anyons n+1, . . . , m, located outside the interferometry region (collectively written as C); and N probe anyons, which we will not number (each of these is written as Bj). We assume the probe anyons initially have no entanglement with the anyons 1, . . . , m, nor with each other. This means that each probe anyon really has at least one other anyon with which it is initially entangled, giving a fourth category of anyons that we neglect because they will not interact with anything and are simply traced out.

The initial state of each probe anyon is described by a density matrix, which can effectively be written as

where p_(b) is the probability that the probe has topological charge b.

The collective state of anyons 1, . . . , m is described by the density matrix ρ^(A), which can be written (using F-moves if necessary) as a linear combination of the basis elements

In these basis elements, a_(in) is the collective charge in the ket of anyons 1, . . . , n, which are in the interior region of the interferometer, and a_(ex) is the collective charge in the ket of anyons n+1, . . . , m, which are in the exterior region of the interferometer. The charges in the respective bras are denoted with primes (since they can take different values), with the exception of the charges a_(j) localized on the individual anyons, which must have definite values. We emphasize that the density matrix ρ^(A) can have anyons 1, . . . , m in an overall collective charge f, which represents “hidden” anyons that are not being kept track of with which these have anyonic charge entanglement. This means superpositions off can only be incoherent (i.e., classical probabilities). We also note that if this overall charge is trivial f=0, then a_(ex)=ā_(in) and a′_(ex)=a′ _(in).

To obtain the transformation of the density matrix ρ^(A) after a single probe measurement, one performs the following series of operations: tensor the density matrix of the probe with ρ^(A), apply the unitary evolution operator U corresponding to the probe passing through the interferometer, apply the projection operator (and renormalize) corresponding to which outcome was observed by the detectors, and finally trace out the probe (representing the fact that the probe is “discarded” after the measurement and no longer kept track of). For the measurement in which the probe is observed at detector s, this is described by

ρ A ↦ B  [ Π s  U  ( ρ A ⊗ ρ B )  U †  Π s ]  [ Π s  U  ( ρ A ⊗ ρ B )  U †  Π s ] . ( 2.12 )

This gives a POVM measurement of ρ^(A) which has the effect of partially projecting the collective charge of anyons 1, . . . , n and partially decohering their anionic charge entanglement with anyons n+1, . . . , m. It may be represented by multiplying the basis elements in Eq. (2.11) by appropriate factors. We are not so concerned with the details of the transformation of ρ^(A) that results from sending a single probe through the interferometer, but rather are particularly interested in the effect on the state ρ^(A) after many probes have passed through the interferometer. In the asymptotic limit of sending N→∞ probe anyons through the interferometer, the effect on ρ^(A) may effectively be treated as the simultaneous projection of the collective charge of anyons 1, . . . , n and the decoherence of their anyonic charge entanglement with anyons n+1, . . . , m.

To be more precise, we focus on the part of Eq. (2.11) that encodes the anionic entanglement between the collective charge of anyons 1, . . . , n and the collective charge of anyons n+1, . . . , m, and apply an F-move to it

Multiple anyons are allowed superpositions of their collective charge, so the density matrix will generally be a linear combination of elements like these. Written in this form, the anyonic charge entanglement between the anyons 1, . . . , n in the interior of the interferometer and anyons n+1, . . . , m in the exterior of the interferometer reduced to a particularly clean description in terms of the single charge e line. These clusters of anyons are said to have no anyonic charge entanglement between them if e=0 is the only charge that has non-zero density matrix coefficients.

One now must determine what charges the probe is capable of distinguishing by monodromy (the full 2π winding of one anyon around another). For this, we consider the expectation value

$\begin{matrix} {M_{aB} = {\sum\limits_{b}\; {p_{b}M_{ab}}}} & (2.15) \end{matrix}$

of the scalar component of the monodromy of the charges a and b

$\begin{matrix} {M_{ab} = {\frac{\left\lbrack {R_{ba}R_{ab}} \right\rbrack}{_{ab}} = {\frac{S_{ab}S_{00}}{S_{0\; a}S_{0b}}.}}} & (2.16) \end{matrix}$

We use this to sort the anyonic charges into classes that the probe anyons can distinguish via monodromy, i.e., all charges a with the same M_(aB) as a particular a_(κ) compose the class

C _(κ) ={aεC:M _(aB) =M _(a) _(κ) _(B)}.  (2.17)

We define the projection operator onto the charge class C_(κ) as

$\begin{matrix} {\Pi_{C_{\kappa}} = {\sum\limits_{c \in C_{\kappa}}\; {\Pi_{c}.}}} & (2.18) \end{matrix}$

The effect of interferometry measurement in the asymptotic limit will observe the charge class C_(κ) with probability

Prob_(A)(κ)={tilde over (T)}r[π _(C) _(κ) ^((1 . . . n))ρ^(A)],  (2.19)

for which it will transform the state ρ^(A) in a way that may be described by applying the following transformation to the basis elements

To elucidate this result and highlight how it differs from projective measurement, we can split the effect into the two parts: projective measurement and decoherence of anyonic charge entanglement. A projective measurement of the collective charge of the anyons 1, . . . , n that observes the charge class C_(κ) [i.e., application of Eqs. (2.3,2.4) using π_(C) _(κ) ^((1 . . . n))] transforms ρ^(A) in a way that may be described by

The decoherence of the anyonic charge entanglement between the collective charge of anyons 1, . . . , n and that of anyons n+1, . . . , m is described by the superoperator D_(B) ^(1 . . . n). The transformation ρ^(A)

D_(B) ^(1 . . . n)(ρ^(A)) may be described by the basis element transformation

This decoherence results from probe anyons passing between the anyon clusters in the interior and exterior of the interferometer, and exhibits an effect specific to non-Abelian anyons. In particular, while one expects off-diagonal a_(in), a′_(in) terms of the density matrix to decohere when the probe can distinguish between a_(in) and a′_(in) (a familiar effect in quantum mechanics), there are also terms that will decohere even when the probe cannot distinguish between a_(in) and a′_(in) (potentially even for a_(in)=a′_(in)). Specifically, this decoherence occurs for terms corresponding to e-channels that the probe can “see” (i.e., M_(eB)≠1), which can exist for a_(in) and a′_(in) that the probe cannot distinguish only for non-Abelian anyons, because they have multiple fusion channels. The decoherence superoperator D_(B) ^((1 . . . n)) obviously has no effect on a state if all the e that occur in this state are such that M_(eB)=1. In particular, under such conditions, an interferometry measurement (in the asymptotic limit) is a projective measurement. This is always the case when the charge measurement result is a definite Abelian charge c. We also emphasize that this decoherence (and thus interferometry measurements) can change the overall collective charge of the anyons 1, . . . , m. In other words, by applying an F-move to return to the original basis, one finds that it is not generally true that

$\begin{matrix} {{\sum\limits_{e}{\left\lbrack \left( F_{a_{i\; n}^{\prime}a_{ex}^{\prime}}^{a_{i\; n}a_{ex}} \right)^{- 1} \right\rbrack_{fe}{D_{e,B}\left\lbrack F_{a_{i\; n}^{\prime}a_{ex}^{\prime}}^{a_{i\; n}a_{ex}} \right\rbrack}_{{ef}^{\prime}}}} = {\delta_{{ff}^{\prime}}.}} & (2.26) \end{matrix}$

Combining these, we find that an interferometry measurement of anyons 1, . . . , n using N→∞ probes B resulting in the outcome C_(κ) is given by

Prob_(A)(κ)={tilde over (T)}r[π _(C) _(κ) ^((1 . . . n))ρ_(A)]  (2.27)

ρ^(A)

Δ_(B;κ) ^((1 . . . n))[ρ^(A) ]=D _(B) ^((1 . . . n))∘π_(C) _(κ) ^((1 . . . n))[ρ^(A)].  (2.28)

Note that the transformations π_(C) _(κ) and D_(B) commute with each other, so the order of their composition is unimportant. If M_(eB)=1 for all e that occur for a_(in), a′_(in) εC_(κ) with non-zero density matrix coefficients, then the action of D_(B) is trivial and interferometry measurement is simply a projective measurement.

If the collective anyonic charge a_(in), a′_(in) is Abelian, then e=0 is the only possible fusion channel, and the superoperator D_(B) ^((1 . . . n)) has no effect, since M_(0b)=1 for all b.

It is often the case that the probe anyons used for interferometry are such that all charges are distinguishable by the probe, i.e., making all charge classes C_(κ) singletons, and only vacuum has trivial monodromy with the probe, i.e., M_(aB)=1 iff a=0. This is the case for all examples considered in this paper, so we restrict our attention to this case from now on. When this is the case, the intfeferometry measurement transformation Δ_(B,c) ^((1 . . . n)) simplifies so that the projection is onto a definite charge (π_(c) ^((1 . . . n))) and anyonic charge entanglement between anyons 1, . . . , n and anyons n+1, . . . , m is completely decohered (D_(e,B)=δ_(0e)), i.e., for a charge measurement outcome c this is represented on the basis elements by

This form makes it particularly easy to apply the interferometry measurement to a state, as one only needs to re-write the diagrams (using F-moves) describing the state to take the form of the left hand side of this equation and then apply this simple transformation.

A point that is glossed over in the above discussion is that there is a phase parameter β in the interference term of the outcome probabilities that is an experimental variable. For example, the tunneling probability for a double-point contact interferometer in the weak tunneling limit is

p _(a) _(in) ^(←˜|t) ₁|² +|t ₂|²+2|t ₁ t ₂ |Re{M _(a) _(in) _(B) e ^(iβ)}  (2.30)

where t_(i) and t₂ are the tunneling amplitudes of the two point contacts. The interferometer can only distinguish between target charges a_(in) and a′_(in) if they give different values of this tunneling probability p_(a) _(in) ^(←)≠p_(a′) _(in) ^(←), which leads to out definition of C_(κ). There are two values (mod 2π) of β which give p_(a) _(in) ^(←)≠p_(a′) _(in) ^(←) when M_(a) _(in) _(B)≠M_(a′) _(in) _(B), but these are non-generic experimental variable values that require the ability to tune β with infinite precision, which is, of course, physically impossible. While we do not need to worry that these non-generic points will make charges indistinguishable, it is important to stay as far away from them as possible in order to increase the distinguishability of the measured charge, which decreases the number probes needed for a given level of confidence in the measurement outcome. Specifically, to distinguish between target charges a_(in) and a′_(in) with a confidence level 1−α, one needs

$\begin{matrix} {{N \succsim {{2\left\lbrack {{erfc}^{- 1}(\alpha)} \right\rbrack}^{2}\left\lbrack \frac{\sqrt{p_{a_{in}}^{\leftarrow}\left( {1 - p_{a_{in}}^{\leftarrow}} \right)} + \sqrt{p_{a_{in}^{\prime}}^{\leftarrow}\left( {1 - p_{a_{in}^{\prime}}^{\leftarrow}} \right)}}{\Delta \; p} \right\rbrack}^{2}},{where}} & (2.31) \\ {{\Delta \; p} = {{{p_{a_{in}}^{\leftarrow} - p_{a_{in}^{\prime}}^{\leftarrow}}}.}} & (2.32) \end{matrix}$

In order to minimize the necessary duration of a measurement, one should tune β so as to maximize Δp. For the double-point contact interferometer in the weak tunneling limit, this takes the maximal value

Δp _(max)=2|t ₁ t ₂ |ΔM,  (2.33)

where ΔM=|M_(a) _(in) _(B)−M_(a′) _(in) _(B)| generally determines how distinguishable a_(in) and a′_(in) are by B probes. For such an appropriately tuned interferometer with |t₁|˜|t₂|˜t<¼, we have the estimate

$\begin{matrix} {N \succsim {{8\left\lbrack \frac{{erfc}^{- 1}(\alpha)}{t\; \Delta \; M} \right\rbrack}^{2}.}} & (2.34) \end{matrix}$

Quantum State Teleportation by “Forced Measurement”

Quantum entanglement is the primary source of philosophical quandaries concerning the foundations of quantum physics, and yet it is also the primary resource of quantum information science. One of the abecedarian examples of entanglement's use as a resource is quantum state teleportation, which enjoys the validation of having been experimentally realized. A novel use of teleportation is as a means of incorporating the entanglement needed to perform a quantum computation. This is the concept underlying measurement-based approaches to conventional quantum computation. Not surprisingly, our measurement-only approach to topological quantum computation utilizes an anyonic version of quantum state teleportation to incorporate the necessary entanglement. This is done by using teleportation to generate the braiding transformations that in turn comprise the computational gates. In order to extend the concept of teleportation to anionic states, we introduce and employ a protocol that we call “forced measurement.” The idea is to perform a quantum state teleportation, without the ability to apply unitary operators to qubits, simply by performing a series of measurements until the desired outcome is achieved. Forced measurement is a probabilistically determined adaptive series of measurements in which the measurements to be carried out are predetermined, but the number of times that they need to be carried out is probabilistically determined by the first attainment of the desired measurement outcome. In this section, we will first show how such a forced measurement protocol would be implemented for conventional teleportation; then we apply it to anyons, reviewing its implementation for projective measurement in the density matrix formalism, and finally describing the modified protocol for implementing anyonic teleportation using interferometry measurements.

Conventional Qubits (Spin-1/2 Systems)

We begin by recalling the usual procedure for quantum teleportation of conventional qubits. Let us write the standard Bell states

$\begin{matrix} {{\Phi^{\pm}\rangle} = {\frac{1}{\sqrt{2}}\left( {{\left. \uparrow\uparrow \right.\rangle} \pm {\left. \downarrow\downarrow \right.\rangle}} \right)}} & (3.1) \\ {{\Psi^{\pm}\rangle} = {\frac{1}{\sqrt{2}}\left( {{\left. \uparrow\downarrow \right.\rangle} \pm {\left. \downarrow\uparrow \right.\rangle}} \right)}} & (3.2) \end{matrix}$

in the following way

$\begin{matrix} {{\Phi_{0}\rangle} = {\frac{1}{\sqrt{2}}{\left( {{\left. \uparrow\uparrow \right.\rangle} + {\left. \downarrow\downarrow \right.\rangle}} \right).}}} & (3.3) \\ {{\Phi_{\mu}\rangle} = {{ \otimes \sigma_{\mu}}{\Phi_{0}\rangle}}} & (3.4) \end{matrix}$

where σ₀=

and

_(j) are the Pauli matrices. The (normalized) state to be teleported is

|ψ

=ψ_(↑)|↑

+ψ_(↓)|↓

  (3.5)

The entanglement resource for the teleportation will be a Bell state |Φ₀

, which we create and then place its two spins at locations 1 and 2. The spin system encoding the state Ψ is at location 3, which is in close enough proximity to 2 so that the collective state of spins 2 and 3 can be measured. This gives the tensor product of states

$\begin{matrix} {{{\Phi_{0}\rangle}_{12}\; {\psi\rangle}_{3}} = {\frac{1}{2}{\sum\limits_{{\mu = 0},\; \ldots \;,\; 3}{\sigma_{\mu}^{(1)}{\psi\rangle}_{1}{\Phi_{\mu}\rangle}_{23}}}}} & (3.6) \end{matrix}$

upon which we perform an orthogonal projective measurement in the Bell basis on spins 2 and 3 (i.e., using π_(μ) ⁽²³⁾=|Φ_(μ)

₂₃

Φ_(μ)|₂₃), that the state transforms as

|Φ₀

₁₂|ψ

₃

π_(μ) ⁽²³⁾[|Φ₀

₁₂|ψ

₃]=σ_(μ) ⁽¹⁾|ψ

₁|Φ_(μ)

₂₃  (3.7)

with probability Prob (μ)=1/4. The usual prescription for teleportation at this point would send the measurement outcome μ (two classical bits of information) from someone named “Bob” at position 2 to someone named “Alice” at position 1, who then applies σ_(μ) to the spin there to recover the state |Ψ

₁ (assuming that someone named “Charlie” or “Eve” has not decided to complicate the situation).

There is however an alternate way to obtain the state at position 1 without applying a recovery Pauli transformation. Notice that in addition to teleporting the state Ψ to position 1, the entanglement resource Bell state has also been teleported to positions 2 and 3. As long as the measurement method is non-demolitional, the entanglement resource may be used again. If we now measure the spins at 1 and 2 in the Bell basis, we get outcome ν with probability Prob (ν)=¼, and the state transforms as

σ_(μ) ⁽¹⁾|Ψ

₁|Φ_(μ)

₂₃

π_(μ) ⁽¹²⁾[σ_(μ) ⁽¹⁾|Ψ

₁|Φ_(μ)

₂₃]=±|Φ_(ν)

₁₂σ_(ν) ⁽³⁾|Ψ

₃.  (3.8)

The ±comes from the fact that σ_(μ)σ_(ν)=σ_(ν)σ_(μ) if μ=0,ν=0, or μ=ν and σ_(μ)σ_(ν)=−σ_(ν)σ_(μ), otherwise. We can now try to teleport the state from position 3 to 1 again, obtaining outcome μ₂ with probability Prob (μ₂)=¼ and the state transformation

|Φ_(ν)

₁₂σ_(ν) ⁽³⁾|Ψ

₃

π_(μ2) ⁽²³⁾[|Φ_(ν)

₁₂σ_(ν) ⁽³⁾|Ψ

₃]=±σ_(μ2) ⁽¹⁾|Ψ

₁|Φ_(μ2)

₂₃.  (3.9)

This can be repeated indefinitely, or rather until the n^(th) try when we get the desired measurement outcome μ_(n)=0, which gives us the state |ψ

₁ at position 1 without a Pauli transformation. Each attempt has probability Prob (0)=¼ of obtaining this desired outcome, so the average number of attempts needed to obtain the desired outcome is

n

=4  (3.10)

and the probability of needing n>N attempts to obtain the desired outcome is

$\begin{matrix} {{{{Prob}\mspace{11mu} \left( {\mu_{1},\ldots \mspace{14mu},{\mu_{N} \neq 0}} \right)} = \left( \frac{3}{4} \right)^{N}},} & (3.11) \end{matrix}$

i.e., failure is exponentially suppressed in the number of attempts. This gives a series of measurements with the set of outcomes M={μ_(n)=0, ν_(n), . . . , μ₁, ν₁=0} such that μ_(n)=ν₁=0 the ν₁=0 initialization is included for convenience). We call such a series of measurements a “forced measurement” (because continue measuring until we get the desired outcome), and write its corresponding operator as

{hacek over (π)}_(M) ^((23←12))=π_(μ) _(n) ₌₀ ⁽²³⁾∘π_(ν) _(n) ⁽¹²⁾∘ . . . ∘π_(μ) ₁ ⁽²³⁾∘π_(ν) ₁ ₌₀ ⁽¹²⁾,  (3.12)

so that

{hacek over (π)}_(M) ^((23←12))[|Φ₀

₁₂|Ψ

₃]=(−1)^(M)|Ψ

₁|Φ₀

₂₃,  (3.13)

where the (irrelevant) overall sign depends on the series of measurement outcomes.

In density matrix notation, writing

ρ=|Ψ

Ψ|  (3.14)

ρ^(ER)=|Φ0

Φ0|,  (3.15)

this forced measurement teleportation takes the form

{hacek over (π)}_(M) ^((23←12))[ρ^(ER(12))

ρ⁽³⁾]=ρ⁽¹⁾

ρ^(ER(23)).  (3.16)

In the context of conventional qubits, the “forced measurement” prescription is obviously a much less efficient alternative to the usual recovery procedure (which is presumably why it is never considered), as it requires multiple uses of a quantum channel in order to shuttle spin system 2 back and forth between spins 1 and 3, rather than two uses of a classical channel. However, it serves the purpose of a conceptual template for anyonic teleportation, where one does not have an operation equivalent to application of a Pauli matrix to use for recovery of the state.

Anyons

In anyonic systems, it is important to use the forced measurement approach to teleportation, because undesired measurement outcomes have the effect of introducing non-trivial anyonic charge entanglement between the state system and other anyons. Unlike the Spin-½ case, correcting for undesired anyonic charge entanglement is not as simple as applying a local operator to one of the anyons, so there is no longer a more efficient alternative to forced measurement. In the case of projective topological charge measurement, the unwanted entanglement introduced by measurement is with the entanglement resource pair. This case is very similar to that of conventional qubits, and essentially the same protocol is used for forced measurement teleportation. In the case of interferometry topological charge measurement, the unwanted entanglement manifests itself in the introduction of a nontrivial overall topological charge of system receiving the teleported state (the entanglement resource pair is automatically unentangled from the teleported state). This is a result of the decoherence aspect of interferometry measurement discussed above that results from entanglement with probe anyons that are taken to “infinity” and traced out. This difference for interferometry measurements requires an additional topological charge measurement to determine whether the desired outcome is achieved, and thus results in a slightly modified forced measurement protocol for teleportation.

Using Projective Measurements

The projective measurement forced measurement procedure using the state vector formalism has been previously introduced. Here we will review those results using the density matrix formalism in order to provide a more uniform treatment of the two cases and allow a more direct comparison with the interferometry forced measurement.

We consider the state encoded in the non-local internal degrees of freedom of some anyons given by the density matrix ρ. Since we are presently only interested in manipulating one particular anyon which has definite charge a, we represent it as

where “ . . . ” represents all the other anyons that comprise the state ρ, which we leave implicit (i.e., there should really be additional anyonic charge lines emanating from the box). Normalization factors giving {tilde over (T)}r[ρ]=1 are absorbed into the box.

In order to teleport the state information encoded by an anyon of definite charge a to another anyon of definite charge a, we introduce a particle-antiparticle pair produced from vacuum, given by the state

The state in Eq. (3.18) has maximal anyonic entanglement between its two anyons, and is the analog of the maximally entangled Bell states typically used as the entanglement resource in quantum state teleportation of conventional qubits.

We now tensor these together

to obtain the combined state on which we will perform measurements. What we would like to do is perform a measurement of the collective charge of anyons 2 and 3 for which the result is vacuum charge 0.

Introducing the notation

it is clear that ρ₀ ^(ER(12)) ρ₀ ^((3 . . . )) (a, . . . )=ρ^(ER(12)) {circle around (x)}ρ^((3 . . . )) (a, . . . ). Applying a projective topological charge measurement of anyons 2 and 3 to the state ρ_(e) ^(ER(12)) ρ_(e) ^((3 . . . )) (a . . . ), we get the measurement outcome f with probability and post-measurement state

Prob(f)=|[f _(a) ^(aāa)]_(ef)|²  (3.21)

π_(f) ⁽²³⁾[ρ_(e) ^(ER(12))ρ_(c) ^((3 . . . ))(a, . . . )]=ρ_(f) ^(ER(32))ρ_(f) ^((1 . . . ))(a, . . . )  (3.22)

To see this, we use the unitary F-move that transforms between the two fusion bases of the three anyons

(which should make clear where the probability Prob (f) comes from) and apply the projective measurement

If we obtain an undesired measurement outcome f≠0, we can “undo” the measurement by subsequently performing a measurement of anyons 1 and 2, as long as the measurement processes are non-demolitional. If this measurement has outcome e₂, the combined system is put in the post-measurement state ρ_(e) ₂ ^(ER(12)) ρ_(e) ₂ ^((3 . . . ))(a, . . . ). Now we can perform a measurement of anyons 2 and 3 again, with an entirely new chance of obtaining the desired outcome. This procedure may be repeated until we obtain the desired measurement outcome, obtaining a string of measurement outcomes M={e₁, f₁, . . . , e_(n)f_(n)} (including the initialization e₁ for convenience), where e₁=f₁=₀ and f_(j)≠0 for j<n. The probabilities of measurement outcomes e_(j) and f_(j) are respectively

Prob(e _(j))=|[F _(a) ^(aāa)]_(e) _(j) _(f) _(j−1) |²  (3.25)

Prob(f _(j))=|[F _(a) ^(aāa) ]e _(j) f _(j)|².  (3.26)

The probability of obtaining the desired outcome f=0 at the j^(th) measurement attempt in this procedure is

$\begin{matrix} {{{Prob}\; \left( {f_{j} = 0} \right)} = {{\left\lbrack F_{a}^{a\; \overset{\_}{a}\; a} \right\rbrack_{e_{j}0}}^{2} = {N_{a\; \overset{\_}{a}}^{e_{j}}{\frac{d_{e_{j}}}{d_{a}^{2}}.}}}} & (3.27) \end{matrix}$

This probability has the non-zero lower bound

Prob(f _(j)=0)≧d _(a) ⁻²,  (3.28)

since d_(x)≧1 for any x. The average number of attempts until a desired outcome is achieved in a forced measurement is thus

n

≦d _(a) ²,  (3.29)

and the probability of needing n>N attempts to obtain the desired outcome is

Prob(f ₁ , . . . f _(N)≠0)≦(1−d _(a) ⁻²)^(N),  (3.30)

i.e., failure is exponentially suppressed in the number of attempts. Thus, the forced projective measurement given by the probabilistically determined adaptive series of measurements

{hacek over (π)}_(M) ^((32←12))=π_(f) _(n) ₌₀ ⁽²³⁾∘π_(e) _(n) ⁽¹²⁾∘ . . . ∘π_(f) ₁ ⁽²³⁾∘π_(e) ₁ ₌₀ ⁽¹²⁾  (3.31)

enables use to perform anyonic state teleportation

{hacek over (π)}_(M) ^((32←12))[ρ^(ER(12))

ρ^((3 . . . ))(a, . . . )]=ρ^(ER(32))

ρ^((1 . . . ))(a, . . . )  (3.32)

using projective measurements. The notation 32←12 means the entanglement resource was originally encoded in anyons 1 and 2, but after the forced measurement is encoded in anyons 3 and 2. One can also infer from this that the role of anyon 3 in encoding the state ρ has now been transferred to anyon 1. We emphasize that while it is important to perform all the π⁽¹²⁾ measurements in order to teleport the state information, the actual outcomes e_(j) of these measurements are unimportant.

Using Interferometry Measurements

In contrast with projective measurement, interferometrical measurement of topological charge is not quite as simple and requires a density matrix formulation. We will assume the asymptotic limit of N→∞ probe measurements in which interferometry may effectively be treated as a projective measurement of the target anyons' collective charge, together with decoherence of anyonic charge entanglement between the target anyons and those exterior to the interferometry region. We will also assume that the probe anyons can distinguish between all topological charges and only has trivial monodromy with the vacuum charge, allowing us to use Eq. (2.29) for interferometry measurements, which can be described as the combination of projection onto a definite charge (π_(c)) and complete anyonic charge decoherence between the interior and exterior of the interferometry region (D_(e,B)=δ_(0e)). Since we otherwise do not care about the specific details of the probe anyons, we will drop the label B and simply denote interferometry measurements using Δ_(c)=D∘π_(c).

We again consider the state of some anyons given by the density matrix ρ, and are presently only interested in manipulating one particular anyon which has definite charge a, so we represent it as

where “ . . . ” represents all the other anyons that comprise the state ρ, which we leave implicit (i.e., there should really be additional anyonic charge lines emanating from the box). We will need to assume that ρ has trivial overall charge ρ=ρ₀ in order to be able to implement our forced measurement procedure. This is not a severe restriction, as one can always achieve such a condition by keeping track of additional anyons. Furthermore, the anyonic states used to encode topological qubits will have trivial overall charge, so this restriction automatically includes those of interest to us for TQC. One might worry that such a restriction is pointless, since we are going to perform interferometry measurements, and we know that interferometry can change the overall charge. In fact, it is exactly this effect that we correct for by using forced measurement. We will see that the interferometry measurements employed are such that they change the overall charge and introduce anyonic charge entanglement in a very particular way that be represented by the attachment of a charge line connecting the a lines on either end of the density matrix. Thus, we define the density matrix that can occur at intermediate steps in the forced measurement where χ=0 obviously gives the original state ρ₀, and the normalization factor is included so that {tilde over (T)}r [ρ_(χ)]={tilde over (T)}r [ρ₀]=1. We note that measuring the collective charge of all the anyons that comprise this state (i.e. measuring the overall charge of ρ_(χ)) gives the measurement outcome χ. Hence, the restriction that the original state ρ₀ have trivial overall charge makes it possible to correctly identify the charge χ by performing a measurement¹. We will later see that it is essential in our forced measurement procedure to be able to properly identify the charge χ, hence the reason for restricting to such initial ρ₀.

The entanglement resource for interferometry measurement based teleportation is

We do not require e=0 here because it is not necessary for the procedure to work (nor does it significantly increase the probability of obtaining a desired measurement result). Furthermore, e will generally change in each step, so if we wanted to require e=0, we would have to perform additional steps to re-initialize the entanglement resource, making the procedure less efficient.

We now tensor these together

Performing an interferometry measurement on anyons 2 and 3 will have the charge measurement outcome f with probability and post-measurement state given by

$\begin{matrix} {\mspace{79mu} {{{Prob}\mspace{11mu} (f)} = {N_{a\; \overset{\_}{a}}^{f}\frac{d_{f}}{d_{a}^{2}}}}} & (3.37) \\ {{\Delta_{f}^{(23)}\left\lbrack {\rho_{e}^{{ER}{(12)}} \otimes {\rho_{x}^{({3\; \ldots}\;)}\left( {a,\ldots}\mspace{14mu} \right)}} \right\rbrack} = {\sum\limits_{y}{{{Prob}\left( {yf} \right)}{\rho_{f}^{{ER}{(32)}} \otimes {\rho_{y}^{({1\; \ldots}\;)}\left( {a,\ldots}\mspace{14mu} \right)}}}}} & (3.38) \\ {\mspace{79mu} {where}} & \; \\ {\mspace{79mu} {{{Prob}\; \left( {yf} \right)} = {\sum\limits_{q}{{\left\lbrack F_{a}^{e\; \overset{\_}{f}\; a} \right\rbrack_{qa}}^{2}\mspace{11mu} {{\left\lbrack F_{a}^{q\; x\; a} \right\rbrack_{y\; a}}^{2}.}}}}} & (3.39) \end{matrix}$

The probability Prob (f) should be obvious, since anyons 2 and 3 are initially unentangled (no charge lines connect them) and thus have randomly determined collective charge (i.e., probabilities weighted by their quantum dimension). To demonstrate the result for the post-measurement state, we break the interferometry measurement into the two parts (projection and decoherence) Δ_(f) ²³=D⁽²³⁾ ∘π_(f) ⁽²³⁾

where the F-move applied in the last step is in preparation for applying the decoherence operator D⁽²³⁾, which then simply projects onto w=0. Applying D⁽²³⁾ to the result of Eq. (3.40) gives the post-interferometry measurement state (using isotopy and braidings whose effects cancel each other) which gives our claimed post-interferometry measurement state. In the last step, we used the diagrammatic evaluation

After the interferometry measurement Δ_(f) ⁽²³⁾, the entanglement resource has moved to anyons 2 and 3 and has no entanglement with the anyons encoding the state ρ₀. However, we are left with incoherent superpositions of the charge y, and the possibility of having an undesired non-trivial charge. We now measure this charge y to determine whether or not we have achieved the desired result y=0. This is done by performing an interferometry measurement on anyons 1, . . . , where “ . . . ” stands for all other anyons that have non-trivial anyonic charge entanglement with anyon 1, which of course now excludes anyons 2 and 3. This measurement will have outcome z (given the previous measurement outcome f) with probability Prob (z|f), and result in the post-measurement state

$\begin{matrix} {{\Delta_{z}^{({1\mspace{14mu} \ldots}\mspace{14mu})}\left\lbrack {\sum\limits_{y}{{{Prob}\left( {yf} \right)}{\rho_{y}^{({1\mspace{14mu} \ldots}\mspace{14mu})}\left( {a,\ldots}\mspace{14mu} \right)}}} \right\rbrack} = {{\rho_{z}^{({1\mspace{14mu} \ldots}\mspace{14mu})}\left( {a,\ldots}\mspace{14mu} \right)}.}} & (3.43) \end{matrix}$

Since the overall charge y was already fully decohered, this interferometry measurement is simply a projective measurement of the incoherent superpositions of y (i.e., this is essentially just a classical measurement).

Thus, the total probability of ending up with ρ_(z) ^((1 . . . )) after performing these two measurement (Δ⁽²³⁾ and Δ^((1 . . . ))) on the state ρ_(e) ^(ER(12)){circle around (x)}ρ_(χ) ^((3 . . . ))(a, . . . ) is

$\begin{matrix} \begin{matrix} {{{Prob}(z)} = {\sum\limits_{f}{{{Prob}\left( {zf} \right)}{{Prob}(f)}}}} \\ {= {\sum\limits_{f,q}{N_{a\overset{\_}{a}}^{f}\frac{d_{f}}{d_{a}^{2}}{\left\lbrack F^{e\overset{\_}{f}a} \right\rbrack_{qa}}^{2}{\left\lbrack F_{a}^{qxa} \right\rbrack_{za}}^{2}(3.45)}}} \end{matrix} & (3.44) \end{matrix}$

and in particular, the probability of achieving the desired outcome z≠0 from this process is

Prob(z=0)=d _(a) ⁻²  (3.46)

To obtain this result, we combined the relation

$\begin{matrix} {{{\left\lbrack F_{a}^{e\overset{\_}{f}a} \right\rbrack_{\overset{\_}{x}a}} = {\sqrt{\frac{f}{d_{f}}}{\left\lbrack F_{a}^{xea} \right\rbrack_{fa}}}},} & (3.47) \end{matrix}$

which can be derived from the pentagon equations, with unitarity of the F-moves to perform the sum

$\begin{matrix} {{\sum\limits_{f}{N_{a\overset{\_}{a}}^{f}\frac{d_{f}}{d_{x}}{\left\lbrack F_{a}^{e\overset{\_}{f}a} \right\rbrack_{\overset{\_}{x}a}}^{2}}} = {{\sum\limits_{f}{N_{a\overset{\_}{a}}^{f}{\left\lbrack F_{a}^{xea} \right\rbrack_{fa}}^{2}}} = 1.}} & (3.48) \end{matrix}$

If the measurement results in an undesired outcome z≠0, we now undo the “failed” teleportation by performing an interferometry measurement Δ⁽¹²⁾ on anyons 1 and 2. For measurement outcome e₂, this gives the post-measurement state

$\begin{matrix} {{\Delta_{e_{2}}^{(12)}\left\lbrack {\rho_{f}^{{ER}{(32)}} \otimes {\rho_{z}^{({1\mspace{14mu} \ldots}\mspace{14mu})}\left( {a,\ldots}\mspace{14mu} \right)}} \right\rbrack} = {\sum\limits_{x_{2}}{{{Prob}\left( {x_{2}e_{2}} \right)}{\rho_{e_{2}}^{{ER}{(12)}} \otimes {\rho_{x_{2}}^{({3\mspace{14mu} \ldots}\mspace{14mu})}\left( {a,\ldots}\mspace{14mu} \right)}}}}} & (3.49) \end{matrix}$

One could now measure χ₂, but this provides no important information and has no significant effect on the forced measurement process, so we will not do so in order to be more efficient with our measurements (especially the more difficult ones involving more than two anyons). Leaving the charge χ₂ unmeasured will simply introduce a sum over the probabilities of having charge χ₂ in the next step. Now we can repeat the measurements Δ⁽²³⁾ and Δ^((1 . . . )) with a new chance of obtaining the desired outcome z₂=0. This procedure may be repeated until the desired outcome is achieved on the nth attempt, giving the set of measurement outcomes

M={e ₁ ,f ₁ ,z ₁ , . . . ,e _(n) ,f _(n) ,z _(n)}  (3.50)

where z_(n)=0 and z_(j)≠0 for j<n (and the initialization e1 is included for convenience). The probabilities of measurement outcomes e_(j), f_(j), and z_(j) are respectively

$\begin{matrix} {{{Prob}\left( e_{j} \right)} = {N_{a\overset{\_}{a}}^{e_{j}}\frac{d_{e_{j}}}{d_{a}^{2}}}} & (3.51) \\ {{{Prob}\left( f_{j} \right)} = {N_{a\overset{\_}{a}}^{f_{j}}\frac{d_{f_{j}}}{d_{a}^{2}}}} & (3.52) \\ \begin{matrix} {{{Prob}\left( z_{j} \right)} = {\sum\limits_{f_{j}}{{{Prob}\left( {z_{j}f_{j}} \right)}{{Prob}\left( f_{j} \right)}}}} \\ {= {\sum\limits_{f_{j},q}{N_{a\overset{\_}{a}}^{f_{j}}\frac{d_{f_{j}}}{d_{a}^{2}}{\left\lbrack F_{a}^{e_{j},{\overset{\_}{f}}_{j},a} \right\rbrack_{qa}}^{2}{{\left\lbrack F_{a}^{{qx}_{j}a} \right\rbrack_{z_{j}a}}^{2}.}}}} \end{matrix} & (3.53) \end{matrix}$

Of course, the only probability that matters for our purposes is the total probability Prob (z_(j)=0)=d_(a) ⁻² from Eq. (3.46) of achieving the desired outcome z_(j)=0 in each attempt, which is independent of j. From this, the average number of attempts n until a desired z_(n)=0 outcome is achieved in an interferometry forced measurement is found to be

n

=d _(a) ²,  (3.54)

and the probability of needing n>N attempts to obtain the desired outcome is exponentially suppressed

Prob(z ₁ , . . . ,z _(N)≠0)=(1−d _(a) ⁻²)^(N),  (3.55)

just as in the previous cases of forced measurement teleportation. For the non-Abelian anyon models considered herein

$\begin{matrix} \begin{matrix} {{{Prob}\left( {z_{j} = 0} \right)} = {\frac{1}{d_{a}^{2}} = \frac{1}{2}}} & {{{for}\mspace{14mu} {Ising}\mspace{14mu} {using}\mspace{14mu} a} = {\frac{1}{2} = \sigma}} \\ {{{Prob}\left( {z_{j} = 0} \right)} = {\frac{1}{d_{a}^{2}} = {\varphi^{- 2} \approx {.38}}}} & {{{for}\mspace{14mu} {Fib}\mspace{14mu} {using}\mspace{14mu} a} = {1 = ɛ}} \\ {{{Prob}\left( {z_{j} = 0} \right)} = {\frac{1}{d_{a}^{2}} = \frac{1}{4{\cos^{2}\left( \frac{\pi}{k + 2} \right)}}}} & {{{for}\mspace{14mu} {{SU}(2)}_{k}\mspace{14mu} {using}\mspace{14mu} a} = \frac{1}{2}} \end{matrix} & (3.56) \end{matrix}$

Thus, the forced interferometry measurement given by the probabilistically determined adaptive series of interferometry measurements

{hacek over (Δ)}_(M) ^((32←12))=Δ_(z) _(n) ₌₀ ^((1 . . . ))∘Δ_(f) _(n) ⁽²³⁾∘Δ_(e) _(n) ⁽¹²⁾∘ . . . ∘Δ_(z) ₁ ^((1 . . . ))∘Δ_(f) ₁ ⁽²³⁾∘Δ_(e) ₁ ⁽¹²⁾  (3.57)

enables us to perform anyonic state teleportation

{hacek over (Δ)}_(M) ^((32←12))[ρ_(e) ₁ ^(ER(12))

ρ₀ ^((3 . . . ))(a, . . . )]=ρ_(f) _(n) ^(ER(32))

ρ₀ ^((1 . . . ))(a, . . . )  (3.58)

using interferometry measurements. We emphasize that while it is important to perform all the Δ⁽¹²⁾ and Δ⁽²³⁾ measurements in order to teleport the state information, the actual outcomes e_(j) and f_(j) of these measurements are unimportant.

It is sometimes possible to use knowledge of measurement outcomes previous to a y_(j) measurement (usually at or near the beginning of a forced measurement) to infer (by classical information processing) that the probability of a measurement with outcome z_(j)=0 is either 0 or 1, thus making the measurement unnecessary (e.g., if any two of e, x, and fare known to be Abelian). This could help with the efficiency of measurement use, but only some fraction of the time.

Measurement-Generated Braiding Transformations

As motivation for this section (and in fact for the entire disclosure), we consider the state of a number anyons, two of which have charge a. Focusing only on these two anyons, we write the state as

If we introduce a charge conjugate pair

nearby (tensor these states together), then applying three projective topological charge measurements, all with outcomes 0, we have where θ_(a)=e^(isπs) ^(a) is a phase, known as the topological spin of a, obtained from straightening out a counterclockwise twist in a charge a line, The dashed line partitions have been inserted to help clarify the contributions to the diagram from each of the three projectors. From this we see that topological charge measurement could potentially be used to generate a braiding transformation. Specifically, performing these three measurements with vacuum 0 as their charge measurement outcomes transforms the state Ψ exactly the same way it would if anyons 1 and 4 were exchanged by a counterclockwise half twist (i.e., braided), up to an overall phase.

However, we know that measurement outcomes are probabilistic, so the only way we could achieve such a transformation via topological charge measurements is if we had some way to force each measurement to have the desired vacuum charge outcome. Of course, this is exactly what we have produced in the Section 3, where we demonstrated that one could perform teleportation through a “forced measurement” procedure.

Based on this observation, we can now produce braiding transformations

for two anyons of definite charge a by introducing an appropriate entanglement resource near these anyons and performing three consecutive forced measurement teleportations.

We assume these two anyons (partially) comprise the state

Using projective topological charge measurements, we introduce the entanglement resource

ρ^(ER) =|ā,a;0

ā,a;0|  (4.6)

and perform three forced projective measurements to get where the forced projective measurements used here are given by

{hacek over (π)}_(M) ₁ ^((21←23))=π_(fn) ₁ ₌₀ ⁽²¹⁾∘π_(en) ₁ ⁽²³⁾∘ . . . ∘π_(f1) ₁ ⁽²¹⁾∘π_(e1) ₁ ₌₀ ⁽²³⁾  (4.8)

{hacek over (π)}_(M) ₂ ^((24←21))=π_(fn) ₂ ₌₀ ⁽²⁴⁾∘π_(en) ₂ ⁽²¹⁾∘ . . . ∘π_(f1) ₂ ⁽²⁴⁾∘π_(e1) ₂ ₌₀ ⁽²¹⁾  (4.9)

{hacek over (π)}_(M) _(a) ^((23←24))=π_(f) _(ng) ₌₀ ⁽²³⁾∘π_(e) _(ng) ⁽²⁴⁾∘ . . . ∘π_(f1) ₃ ⁽²³⁾∘π_(e1) ₃ ₌₀ ⁽²⁴⁾.  (4.10)

In these series, e₁ ₁ =0 by the initialization assumption, e₁ ₂ =f_(n) ₁ , and e₁ ₃ =f_(n) ₂ , so the first measurement of each forced measurement series is actually already done by the previous forced measurement series and hence need not be repeated, Similarly, we have

{hacek over (π)}_(M) ₃ ^((23←21))∘{hacek over (π)}_(M) ₂ ^((21←24))∘{hacek over (π)}_(M) ₁ ^((24←23))[ρ^(ER(23))

ρ^((14 . . . ))(a,a, . . . )]=ρ^(ER(23))

R _(aa) ^(\(14))ρ^((14 . . . ))(a,a, . . . )R _(aa) ⁽¹⁴⁾.  (4.11)

using the forced projective measurements

{hacek over (π)}_(M) ₁ ^((24←23))=π_(fn) ₁ ₌₀ ⁽²⁴⁾∘π_(en) ₁ ⁽²³⁾∘ . . . ∘π_(f1) ₁ ⁽⁴²⁾∘π_(e1) ₁ ₌₀ ⁽²³⁾  (4.12)

{hacek over (π)}_(M) ₂ ^((21←24))=π_(fn) ₂ ₌₀ ⁽²¹⁾∘π_(en) ₂ ⁽²⁴⁾∘ . . . ∘π_(f1) ₂ ⁽²¹⁾∘π_(e) ₁₂ ₌₀ ⁽²⁴⁾  (4.13)

{hacek over (π)}_(M) ₃ ^((23←21))=π_(fn) ₃ ₌₀ ⁽²³⁾∘π_(en) ₃ ⁽²¹⁾∘ . . . ∘_(f1) ₃ ⁽²³⁾∘π_(e1) ₃ ₌₀ ⁽²¹⁾.  (4.14)

Eqs. (4.7,4.11) respectively generate the counterclockwise and clockwise braiding transformations of anyons 1 and 4 on the state ρ. Thus, with the introduction of an appropriate entanglement resource ρ^(ER) pair of anyons at positions 2 and 3, we can schematically write appropriate series of projective topological charge measurements as being equivalent to braiding transformations of anyons of charge a at positions 1 and 4

{hacek over (π)}_(M) ₃ ^((23←24))∘{hacek over (π)}_(M) ₂ ^((24←21))∘{hacek over (π)}_(M) ₁ ^((21←23)) ≅R _(aa) ⁽¹⁴⁾  (4.15)

{hacek over (π)}_(M) ₃ ^((23←21))∘{hacek over (π)}_(M) ₂ ^((21←24))∘{hacek over (π)}_(M) ₁ ^((24←23)) ≅R _(aa) ^(\(14))  (4.16)

When using interferometry measurements of topological charge, we restrict the state ρ(a, a, . . . )=ρ₀(a, a, . . . ) to have overall charge 0 (because this restriction was needed to perform forced interferometry measurement anyonic teleportation), and introduce the entanglement resource

$\begin{matrix} {\rho_{e}^{ER} = {\frac{1}{d_{e}}{{\overset{\_}{a},{a;e}}\rangle}\mspace{11mu} {\langle{\overset{\_}{a},{a;e}}}}} & (4.17) \end{matrix}$

where the particular value of e is unimportant. To generate the braiding transformations, we perform three forced interferometry measurements where e′=f_(n) ₃ is the collective charge (whose value is unimportant) of the entanglement resource pair given by the last f charge measurement from the third forced measurement series M₃. The forced interferometry measurements used here are given by

{hacek over (Δ)}_(M) ₁ ^((21←23))=Δ_(zn) ₁ ₌₀ ^((34 . . . ))∘Δ_(fn) ₁ ⁽²¹⁾∘Δ_(en) ₁ ⁽²³⁾∘ . . . ∘Δ_(z1) ₁ ^((34 . . . ))∘Δ_(f1) ₁ ⁽²¹⁾∘Δ_(e1) ₁ ⁽²³⁾  (4.19)

{hacek over (Δ)}_(M) ₂ ^((24←21))=Δ_(zn) ₂ ₌₀ ^((13 . . . ))∘Δ_(fn) ₂ ⁽²⁴⁾∘Δ_(en) ₂ ⁽²¹⁾∘ . . . ∘Δ_(z1) ₂ ^((13 . . . ))∘Δ_(f1) ₂ ⁽²⁴⁾∘Δ_(e1) ₂ ⁽²¹⁾  (4.20)

{hacek over (Δ)}_(M) ₂ ^((23←24))=Δ_(zn) ₃ ₌₀ ^((14 . . . ))∘Δ_(fn) ₃ ⁽²³⁾∘Δ_(en) ₃ ⁽²⁴⁾∘ . . . ∘Δ_(z1) ₂ ^((14 . . . ))∘Δ_(f1) ₂ ⁽²³⁾∘Δ_(e1) ₂ ⁽²⁴⁾,  (4.21)

where e₁ ₂ =f_(n) ₁ and e₁ ₃ =f_(n) ₂ . Again, the first measurement of each forced measurement series is actually already done by the previous forced measurement series, so they need not be repeated. Similarly, we have

{hacek over (Δ)}_(M) ₃ ^((23←21))∘{hacek over (Δ)}_(M) ₂ ^((21←24))∘{hacek over (Δ)}_(M) ₁ ^((24←23))[ρ_(e) ^(ER(23))

ρ₀ ^((14 . . . ))(a,a, . . . )]=ρ_(e′) ^(ER(23))

R _(aa) ^(\(14))ρ₀ ^((14 . . . ))(a,a, . . . )R _(aa) ⁽¹⁴⁾,  (4.22)

using the forced interferometry measurements

{hacek over (Δ)}_(M) ₁ ^((24←23))=Δ_(zn) ₁ ₌₀ ^((13 . . . ))∘Δ_(fn) ₁ ⁽²⁴⁾∘Δ_(en) ₁ ⁽²³⁾∘ . . . ∘Δ_(z1) ₁ ^((13 . . . ))∘Δ_(f1) ₁ ⁽²⁴⁾∘Δ_(e1) ₁ ⁽²³⁾  (4.23)

{hacek over (Δ)}_(M) ₂ ^((21←24))=Δ_(zn) ₂ ₌₀ ^((34 . . . ))∘Δ_(fn) ₂ ⁽²¹⁾∘Δ_(en) ₂ ⁽²⁴⁾∘ . . . ∘Δ_(z1) ₂ ^((34 . . . ))∘Δ_(f1) ₂ ⁽²¹⁾∘Δ_(e1) ₂ ⁽²⁴⁾  (4.24)

{hacek over (Δ)}_(M) ₃ ^((23←21))=Δ_(zn) ₃ ₌₀ ^((14 . . . ))∘Δ_(fn) ₃ ⁽²³⁾∘Δ_(en) ₃ ⁽²¹⁾∘ . . . ∘Δ_(z1) ₃ ^((14 . . . ))∘Δ_(f1) ₃ ⁽²³⁾∘Δ_(e1) ₃ ⁽²¹⁾.  (4.25)

Eqs. (4.18,4.22) respectively generate the counterclockwise and clockwise braiding transformations of anyons 1 and 4 on the state ρ₀. Thus, with the introduction of an appropriate entanglement resource ρ_(e) ^(ER) (with arbitrary e allowed) pair of anyons at positions 2 and 3, we can schematically write appropriate series of interferometry topological charge measurements as being equivalent to braiding transformations of anyons of charge a at positions 1 and 4

{hacek over (Δ)}_(M) ₃ ^((23←24))∘{hacek over (Δ)}_(M) ₂ ^((24←21))∘{hacek over (Δ)}_(M) ₁ ^((21←23)) ≅R _(aa) ⁽¹⁴⁾  (4.26)

{hacek over (Δ)}_(M) ₃ ^((23←21))∘{hacek over (Δ)}_(M) ₂ ^((21←24))∘{hacek over (Δ)}_(M) ₁ ^((24←23)) ≅R _(aa) ^(\(14)).  (4.27)

An important point to emphasize is that, for both versions of measurement-generated braiding, the entanglement resource is fully replenished and returned to its original location at the end of these processes. This allows such measurement-generated braiding transformations to be employed repeatedly, without exhausting the entanglement resources. This is an advantage over the measurement-based approaches to conventional quantum computation, for which the entanglement resources are consumed by the computation processes.

In writing the particular measurements and corresponding diagrams in this section, there are implicit assumptions of the relative spatial configurations of the anyons and the manner in which the measurements are performed. It is important to understand that the precise positions and measurement geometries are not important, but rather it is the topologically invariant quantities that determine how the corresponding diagrams should be written. This can efficiently be encapsulated by the delineation of the spatial region (with respect to the anyon locations) for which the topological charge is being measured. The 2-dimensional spatial configuration of the anyons and projective measurements used to generate the braiding transformation in Eq. (4.15,4.16) are shown in FIG. 2. There is, of course, also some freedom to make topologically different choices in the configurations and measurements used. This will generally change the details of the forced measurement procedures that must be used in a manner dependent upon the details of the configuration choice, however there are no conceptual differences between these procedures. When using interferometry measurements, one must be careful not to allow the probe anyons to cause decoherence of important anyonic charge entanglement in the computational anyons' system. We recall from above that a stream of probe anyons passing between two regions will decohere the anyonic charge entanglement between anyons in the different regions. Thus, we need to make sure that the paths of probe anyons avoid partitioning space into regions that severe important charge entanglement lines in the computational anyons' system. We will therefore also indicate the probes' entry and exit paths (which will be chosen in to be “safe”) when delineating the spatial regions of measurement. One also needs to be a bit careful to properly treat the measurements including “ . . . ” anyons. The 2-dimensional spatial configuration of the anyons and interferometry measurements used to generate the braiding transformation in Eq. (4.26,4.27) are shown in FIG. 3.

Measurement-Only Topological Quantum Computation

Once we know how to generate braiding transformations using only topological charge measurements, it is clear that, given an appropriate array of anyons encoding qubits and providing necessary entanglement resources, we can perform topological quantum computation using topological charge measurements as the only computational primitive. We arrange our initialized computational anyons in a linear array and distribute maximally entangled pairs (more or less) between them, forming a quasi-one-dimensional array, as in FIG. 4. These anyons all remain stationary and computational gates on the topological qubits are implemented via measurements. Any quantum computation algorithm can be written in terms of a set of universal gates, using a quantum circuit model. Topological quantum computation in turn is based on the fact that computational gates, in particular those in the universal gate set one chooses to employ for a computation, can be generated (to arbitrary precision) from a series of braiding transformations of the computational anyons encoding the topological qubits. The relations in Eqs. (4.15, 4.16, 4.26, 4.27) give the maps between braiding transformations and the topological charge measurements that will generate them. Combining these determines the series of measurements that should be performed to implement a particular quantum algorithm. It is important to remember that each forced measurement used in this implementation is a probabilistically determined adaptive series of measurement in the sense that a pre-specified pattern of measurements is repeated until a desired outcome is obtained. Since each repetition of the measurement pattern has a non-zero lower bound probability of resulting in the desired outcome, the number of repetitions needed to complete each forced measurement will be exponentially suppressed. This means that even though the time taken by each forced measurement is probabilistically determined, unacceptable delays (from the perspective of computational time scales) will generally not occur as a result of employing this procedure.

The implementation of measurement-only topological quantum computation described above is completely straightforward for the case of projective measurements, but there is a point one must be careful with when considering interferometry measurements that requires us to consider how qubits are encoded more carefully. There are many ways quantum information can be encoded in the non-local multi-dimensional Hilbert space of multiple non-Abelian anyons, and at some level the specific choice of how to encode is unimportant. However, different encodings have different properties, some of which may prove to be advantageous or detrimental, so in practice, the choice can be very important. For the sake of physically implementing a computational model, we require that there be a fixed upper bound (for arbitrary computation size) on the total number of anyons upon which any single measurement is performed. This is important because when we have a stationary array of computational and entanglement resource anyons, increasing the number of anyons in a measurement increases the perimeter of the measurement area, which in turn will generally decrease the fidelity and quantum coherence of the measurement. For example, the probe anyons in FQH interferometers are edge excitations, which have a finite coherence length determined by the system. For the case of projective MOTQC, this bound is automatically satisfied, as measurements are only performed on pairs of anyons. For interferometrical MOTQC (as we have described it), there are interferometry measurements involving “ . . . ” anyons. A careless choice of encoding could require such measurements to include all the computational anyons, a number that scales with the number of qubits. Another point to keep in mind is that interferometry measurement generated teleportation and braiding (as we described it) required that the state upon which they act be restricted to have trivial overall topological charge 0. At this point we limit our attention to TQC with anyon models that have only self-dual topological charges. This is not a severe limitation as it includes the most physically relevant examples of non-Abelian anyon models: Ising, Fib, and SU(2)_(k). If the computational anyons have self-dual charge a=ā, we can use a more economical distribution of entanglement resource anyons, situating only one anyon from each maximally entangled pair between each adjacent pair of computational anyons, so that the second row of X's in FIG. 4 is not needed, and the anyon playing the role of anyon 3 in Eqs. (4.7, 4.11, 4.18, 4.22) and FIGS. 2 and 3 is instead situated either to the left of anyon 1 or to the right of anyon 4, depending on to which pair of computational anyons a braiding transformation is being applied. For TQC models in which the computational anyons do not all have the same anyonic charge, the same forced measurement anyonic teleportation generated braiding principles may be applied, but a greater number of entanglement resource anyons will be needed.

The standard choice of computational anyons for our example anyon models is

$\begin{matrix} {{a = {\frac{1}{2}\mspace{14mu} {for}\mspace{14mu} {Ising}}}{a = {1\mspace{20mu} {for}\mspace{14mu} {Fib}}}{a = {\frac{1}{2}\mspace{14mu} {for}\mspace{14mu} {{SU}(2)}_{k}}}} & (5.1) \end{matrix}$

which all obey the fusion rule

a×x=0+1  (5.2)

(where 0=I,

${\frac{1}{2} = \sigma},$

and 1=Ψ for Ising; and 0=I and 1=ε for Fib). This naturally suggests the two fusion channels 0 and 1 as the basis in which to encode a topological qubit. For an encoding that grants us all of the necessary properties for implementing interferometrical MOTQC, we choose topological qubits to be encoded in the possible fusion channels of four charge a non-Abelian anyons that have collective charge 0, so the topological qubit basis states and |0

and |1

are for j=0,1. Generally, there will be multiple fusion channels from which to choose c_(j). For our favorite examples, there are exactly two fusion channels, so we use

c ₀=0 and c ₁=1 for Ising, Fib, and SU(2)_(k).  (5.4)

This encoding is natural in the sense that it reproduces the standard qubit tensor product structure in the anyon model, i.e. there are no anyonic charge lines connecting the different topological qubits, and so no anyonic charge entanglement between them. Combining this property with the property that each 4 anyon topological qubit has overall charge 0 allows us to fix an upper bound on the number of anyons that must be included in any single topological charge measurement employed in interferometrical MOTQC. To see this, we recall from above that the state upon which we perform forced interferometry measurements needed to have overall charge 0 and the “ . . . ” anyons were all anyons that originally had nontrivial anyonic charge entanglement with the anyon whose state is being teleported. Thus, as long as the topological qubits remain in the computational subspace of the encoding (i.e., with each qubit 4-tuple of anyons having overall charge 0), the “ . . . ” only includes the other anyons of the 4 encoding the qubit upon which an operation is being performed. Of course, this is generally only the case for single qubit gates. For multi-qubits gates in TQC, even though the topological qubits start and end in the computational subspace, there will generally be intermediate steps in the series of braiding transformations composing the gate during which there is non-trivial anyonic charge entanglement between the anyons of different topological qubits. For such n-qubit gates, the “ . . . ” measurements could require measuring the collective topological charge of up to 4 n anyon in one measurement. Fortunately, from the quantum circuit model, we know that computations may be performed with an upper bound on n. In fact, we can choose gate sets that only include single and 2-qubit gates, setting the upper bound at 8 anyons the must be included in any single topological charge measurement.

We summarize by stating that the concept of MOTQC is essentially to change how braiding transformations in TQC are implemented. Instead of physically transporting computational anyons around each other, one can instead perform a series of topological charge measurements. In particular, this means that the issue of computational universality of anyon models' braiding statistics is still present. A set of operations in computationally universal if they can densely populated the space of unitary transformations in the computational space. The braiding transformations of the Ising/SU(2)₂ anyons are the generators of the Clifford group, and hence are not computationally universal. To perform quantum computation with such anyons, one must find a way to supplement the braiding transformations with additional gates that makes the gate set universal. Two methods of doing this have been proposed so far: one involving non-topologically protected operations that are then “distilled” using topologically protected gates; the other involving dynamical topology change of the topological fluid in which the anyons exist. A nice feature of the Ising/SU(2)₂ anyons for the sake of implementing MOTQC is that the interferometry measurements that are used are actually projective measurement, so one can simply use the projective MOTQC protocol, instead of interferometry MOTQC. The reason for this is that one always measures pairs of σ anyons, so the topological charge measurement outcomes I and Ψ are always Abelian charges, which, as explained above, means the interferometry measurement is projective. On the other hand, it has been shown that the braiding transformations are computationally universal for the Fibonacci anyons and for SU(2)_(k) when k=3 and k≧5.

Implementation in Fractional Quantum Hall Systems

Fractional quantum Hall systems are (currently) the most physically concrete candidates in which to implement topological quantum computing platforms, so we will address the anyonics for MOTQC devices in FQH systems in further detail. The braiding statistics of quasiparticles in the most physically relevant non-Abelian FQH states may be written in terms of an Ising, Fibonacci, or SU(2)_(k) anyon model times an Abelian sector (see Appendix). For the purposes of TQC, one can simply focus on the non-Abelian sector of the anyon model describing quasiparticles in a FQH state, because the Abelian sector only contributes overall phases that are irrelevant for quantum computation. The topological charges of these anyon models are all self-dual, so we can employ the more economical distribution of entanglement resource anyon pairs needed for MOTQC.

An anyon with definite topological charge a in a FQH system can be realized by either a single quasiparticle excitation that carries that topological charge, or by a cluster of quasiparticles that are always treated collectively and have (one way or another) been projected into a state with definite collective topological charge a. One example of how to controllably localize anyons in FQH systems is through the use of anti-dots. By placing a gate on top of the Hall bar that allows the Hall fluid beneath it to be selectively depleted through application of voltage, an anti-dot can be created that localizes a controllable number of quasiparticle excitations. Quasiparticle excitations localized on a single anti-dot are essentially fused together, and so are treated collectively and have definite collective charge (as a result of superselection). It should be clear that the use of antidots are not easily amenable to physically transporting anyons (i.e., entire antidots) around each other, which demonstrates why MOTQC is a desirable technique.

The initialization step for implementing MOTQC in FQH systems involves forming a quasi-linear array, formed by two rows of anyons. One row contains the computational anyons, and has 4 n anyons in order to encode n topological qubits, as explained above. The second row contains 4 n entanglement resource anyons. To initialize this array of anyons for computation, the anyons in each row are divided into 2 n adjacent pairs which are put into the collective topological charge 0 state. This can be carried out by several techniques, one of which is to simply measure the topological charge of each pair and, if it has the wrong value, throw it away and use a new pair. This initializes each qubit in the |0

state, and each entanglement resource pair to have charge 0.

Topological Charge Measurement

As topological charge measurements are the keystone of MOTQC, it is important to address how they will be carried out in FQH implementations. In order to perform topological charge measurements in a manner that is non-demolitional and leaves the measured anyons stationary, the only choice for FQH systems is to use double point-contact interferometers. Such FQH interferometers have been studied in detail. Furthermore, they have been successfully implemented experimentally for Abelian FQH systems, and recently for non-Abelian FQH systems.

The concept of a double-point contact interferometer is to opportunistically use of the edge current as a natural supply of probe anyons. By deforming the Hall fluid edges, for example using gates placed on the Hall bar, one can form a point contact—a constriction in the Hall fluid where two separate edge regions are brought into close enough proximity that edge quasiparticle excitations can tunnel through the Hall fluid from one edge to the other. Creating two such point contacts around a region of Hall fluid using the same edges establishes an interferometry loop around this region that allows the different tunneling paths to give rise to interference effects that can be experimentally observed in the resulting tunneling current (related to the Hall conductance) from one edge to the other. The interference will depend on the total magnetic flux as well as the total collective topological charge enclosed in the interference loop, and running such a double-point contact interferometer will thus measure the collective topological charge contained inside the interference loop. Because of the chirality of FQH systems, these double-point contact interferometers are of the Fabry-Pérot type, allowing multiple passes of probe anyons around the interferometry loop. To suppress such terms higher order terms, we restrict to point contacts that are in the weak tunneling limit. Restricting to this limit also has the important effect of ensuring that quasiparticle tunneling across the point contacts will be dominated by fundamental quasiholes, which are the most RG relevant contributions because they are the excitations with smallest conformal scaling dimension.

Once such an interferometer has collapsed superpositions of collective topological charge inside it to be a_(in), the tunneling current (in the weak tunneling limit) will be proportional to the tunneling probability

p _(a) _(in) ^(←) ˜|t ₁|² +|t ₂|²+2|t ₁ t ₂ |Re{M _(a) _(in) _(B) e ^(i) ^(β) },  (6.1)

where t₁ and t₂ are the tunneling amplitudes of the two point contacts, is due to the background contribution, and M_(a) _(in) _(B) is the monodromy scalar component for which the probe B is the topological charge of the fundamental quasihole. Deviation from the weak tunneling limit (|t_(j)|<<1) will amount to a uniform suppression factor Qε[0,1] of the interference term. In fact, most factors that decrease the interferometer's coherence can be lumped into this Q factor, such as noise, background and path fluctuations, edge-bulk tunneling, thermal effects, and finite coherence lengths of edge excitations.

Fortunately, the effects encoded in the Q factor do not introduce errors, but, rather, they only reduce the visibility of the quantum interference in the experiment, i.e., Δp is multiplied by Q. Of course, this means the number of probes needed for a desired level of confidence in a measurement is multiplied by Q⁻², so the duration of measurements will need to be increased, but a simple multiplicative constant to such time scales such as this will cause no major trouble.

The key challenge for the purposes of MOTQC is to have such interferometers where we need them only during specified time intervals. For this, we envision an array of gates on the Hall bar that allow us to deform the edge in a controlled fashion used to construct and destruct arm-like protrusions of the edge reaching into the bulk of the Hall fluid that establish interferometers where necessary (see FIG. 5). Creating and removing these interferometer arms must be done carefully so as not to disturb the rest of the system, particularly the topological qubits.

Calibration

We notice from Eq. 6.1 that the tunneling probability distinguishes topological charge measurement outcomes a_(in) by the interference term proportional to

Re{M _(a) _(in) _(B) e ^(i) ^(β) }=|M _(a) _(in) _(B)|cos(θ_(a) _(in) _(B)+β),  (6.2)

where β is an experimental variable, and θ_(a) _(in) _(B)=arg {M_(a) _(in) _(B)}. This is still true when there is a Q factor. Trying to distinguish topological charges a and a′ for which the monodromy matrix elements have the same magnitudes |M_(aB)|=|M_(a′B)| can get a bit tricky. Since these monodromies and their resulting tunneling probabilities and currents are only distinguished by their relative phase, in order to correctly identify the topological charge measurement outcome, one must initially calibrate the interferometers, associating fixed values of tunneling current with the different measurement outcome possibilities (tuning β to a preferable value). Additionally, one must have the ability to reproduce tuned values of β with reliable precision. This may be quite difficult to achieve.

When trying to distinguish topological charges a and a′ for which the monodromy matrix elements have the different magnitudes |M_(aB)|≠|M_(a′B)|, there is a more robust method of identifying the measurement outcome. By varying β, one can observe the magnitude of the interference term, and in doing so distinguish between a and a′. For this, the interferometers' calibration only involves associating interference magnitudes with the different measurement outcome possibilities.

Unfortunately, when using Ising anyons for TQC, we have to deal with the first situation, because we need to distinguish between the Ising charges I and using σ probes, which have M_(1σ), =1 and M_(Ψσ)=−1. As a minor concession, these monodromies are at least also maximally distinguishable (ΔM=2), giving them the most amount of leeway one can hope for when precision tuning of β is required. When using Fibonacci anyons for TQC, we fortunately have the benefit of the second situation, because we need to distinguish between the Fib charges I and e using e probes, which have M_(Iε)=1 and M_(εε)=Φ⁻²=−0.38. When using SU(2)_(k) anyons for TQC, we need to distinguish between the charges 0 and 1 using charge ½ probes, which have

$M_{0\frac{1}{2}} = {{1\mspace{14mu} {and}\mspace{14mu} M_{1\frac{1}{2}}} = {1 - {4\; {{\sin^{2}\left( \frac{\pi}{k + 2} \right)}.}}}}$

For k=2, this is exactly the same as the Ising anyons. For k≧3, these monodromies have different magnitudes, so we have the benefit of the second situation described above. However, we also point out that as k increases,

$\left. M_{1\frac{1}{2}}\rightarrow\left. {1\mspace{14mu} {and}\mspace{14mu} \Delta \; M}\rightarrow 0 \right. \right.,$

so the distinguishability also decreases, eventually to the point of making either method incapable of distinguishing the charges in practice. Thus, k=3, which is exactly the same as the Fibonacci anyons, is the optimal case from the perspective of measuring topological charge.

Time Scales

It is difficult to make a sound estimate of the time τ_(r) it takes to re-pattern the edges of the FQH fluid in order to construct/destruct interferometers as desired. We can, however, make a naïve estimate assuming that the bulk Hall fluid can be displaced a distance L_(int) corresponding to the size of the interferometer, at a rate determined by the velocity ν_(e) of electrons at the edge of the Hall fluid. This gives an estimated interferometer construction/destruction time scale of

τ_(r)˜ν_(e) L _(int).  (6.3)

The electron edge velocity in experiments is found to be about ν^(˜)10³ m/s, and we can approximate the interferometer length as L_(int)˜1 μm, giving the time estimate of τ_(r)˜1 ns. It is quite possible this is an overly optimistic estimate and that other rate determining factors will necessitate longer durations, but more accurate time scales for such processes are, in any case, best obtained empirically.

We can estimate the measurement duration τ_(m) one needs to run a properly tuned double point-contact interferometer (in the weak tunneling limit) to achieve a desired confidence level 1−α for the interferometry topological charge measurement by using Eq. (2.34) to give

$\begin{matrix} {{{\tau_{m}\text{∼}\frac{e^{*}t^{2}N}{I_{t}}} \gtrsim \frac{8\; {e^{*}\left\lbrack {{erfc}^{- 1}(\alpha)} \right\rbrack}^{2}}{{I_{t}\left( {\Delta \; M} \right)}^{2}}},} & (6.4) \end{matrix}$

where e* is the electric charge of the tunneling quasiparticles (i.e., the probe anyons of the interferometer), I_(t) is the tunneling current through a single point contact with tunneling amplitude t, and N is the estimated number of probes needed. For example, to achieve α=10⁻⁴ in the ν=5/2 MR state, where the probes are fundamental quasiholes (which have e/4 electric charge and σ Ising topological charge) and the measurements are distinguishing between I and Ψ Ising topological charges, a tunneling current I_(t)˜1 nA typical of point contacts in experiments conducted at ν=5/2 gives the necessary measurement duration estimate τ_(m)˜1 ns. This analysis neglected a possible suppression Q factor, which, as discussed above, will increase the measurement time needed by a factor of Q⁻², but it is difficult to quantify all the contributions to Q, so is also best to obtain it empirically.

Combining these two time estimates with the average number of attempts

n

˜d_(a) ² needed in each forced measurement (which is equal to 2 for the MR state) and the

number of topological charge measurements in each attempt (2 or 3), we estimate the time it takes to implement a single measurement generated braiding transformation R_(aa) as

τ_(R)˜3d _(a) ²(2τ_(r)+τ_(m)),  (6.5)

For the ν=5/2 MR state, this gives a time estimate for implementing a braiding transformation on the order of τ_(R)˜10 ns. The number of braiding transformations needed will depend on the computation being performed and the anyon model of the non-Abelian FQH state being used.

Error Sources

Our measurement-only scheme is strongly dependent on the accuracy of our measurements. Suppose, for instance, that we measure the total topological charge of a pair of σ quasiparticles in the ν=5/2 MR state. Let us further suppose that we perform the measurement for long enough that the state is projected onto I or Ψ topological charge with the desired accuracy. There is still the possibility that if the state of the pair is I then we may misidentify it as Ψ, which would be a fatal error. How could such a calamity occur? Our measurement technique is interferometry, so we must be able to distinguish the two possible conductances which can occur, depending on the collective topological charge of the pair of σs. In other words, we must be able to distinguish between two possible values of the tunneling current through the interferometer. However, the current will invariably be noisy (if not as a result of thermal noise, then at least as a result of quantum noise). Thus, it is important for us to make the noise as small as possible.

Let us suppose that we are at low enough temperatures that we can ignore all sources of noise apart from quantum noise. We can estimate the temperature which must be maintained in order to avoid Johnson-Nyquist noise. At low frequencies, there will be ‘shot noise,’ which can be understood as follows. During a measurement time τ_(m), N_(t) quasiparticles will tunnel across the interferometer, leading to a current I_(t)=e*N_(t)/τ_(m). If these are independent events with a Poisson distribution, as can be shown to be the case by direct computation, then the fluctuations in the current will be ΔI_(t)=e*√{square root over (2N_(t))}/τ_(m)=√{square root over (2e*I_(t)/τ_(m))}. This is often written as ΔI_(t)=√{square root over (2e*I_(t)Δω)}, where ω is the frequency bandwidth of the measurement. Consequently, we can minimize the fluctuations in the current by performing the measurement over as long a time as possible. For the sake of concreteness, let us suppose that we have configured our interferometer so that the conductance through the interferometer is g>0 if the two quasiparticles inside the interferometer fuse to Ψ and the conductance is zero (perfect destructive interference) is the two quasiparticles fuse to I. Suppose the two quasiparticles fuse to Ψ. If an average current It would flow through the interferometer in a measurement of infinite duration then, for a Poisson distribution, there is a probability e^(−τ) ^(m) ^(I) ^(t) ^(Ie) ^(*) that zero current would flow through the interferometer during a measurement time τ_(m). This probability can be made smaller by making either I_(t) or τ_(m) larger. However, we cannot make I_(t) too large as I_(t)=gV and we need e*V<<Δ, where Δ is the energy gap. Hence, the probability of an erroneous measurement is greater than e^(−τ) ^(m) ^(gΔ/(e) ^(*) ⁾ ² . By making τ_(m) large, we can make this error probability small. In principle, we are free to make τ_(m) as large as we like, so long as it remains much smaller than the decoherence time of the qubit, τ_(m)<<τ_(dec). As we discuss below, we expect τ_(dec)˜e^(Δ/T) for temperature T. Thus, by taking

1/Δ<<τ_(m) <<e ^(Δ/T),  (6.6)

we can make the probability of an erroneous measurement very small. For the ν=5/2 state, the currently largest measured values of the gap Δ_(5/2)˜500 mK, together with experimentally accessed temperatures of about T^(˜)10 mK, indicate that desirable measurement times for avoiding errors should be 0.1 ns<<τ_(m)<<10²¹ ns.

Stray excitations are another source of error. In any real device, there will be local potential wells where quasiparticles get trapped. As long as they do not move, they will not be an overly serious problem, as their effect can be neutralized in a number of ways. In the Ising case, each computational σ anyon can actually be a collection of an odd number of quasiparticles, some of which may be on an antidot, while the others are stray excitations localized in the bulk. In order for this identification to be correct, it is important that, when we perform measurements of the combined topological charge of two σ anyons, we always include the stray localized excitations that we are associating with a computational anyon. Thus, when we group some stray excitations with our anti-dots, it makes sense to include in our grouping those stray excitations which are enclosed by the arms of the possible interferometers and to exclude those which are not. If there are an odd number of quasiparticles on an anti-dot and an odd number of stray localized excitations that is grouped with an anti-dot in this way, then we should either modify the interferometer arms so that they enclose an even number of strays, somehow remove one of the strays, or instead localize an even number of quasiparticles on the anti-dot. The same strategy works for Fibonacci anyons, except that in this case we can have either an even or an odd number of quasiparticles that define a computational anyon, so long as the total topological charge of the group (anti-dot+strays) is ε. If it is I, then we proceed with the same methods as in the Ising case to change the number of quasiparticles comprising the computational anyon. For general non-Abelian anyons, essentially the same strategy applies, but there may be more fusion channels to worry about.

If the stray excitations move around, this is a more serious problem. Motion within a grouping is not a problem since this cannot change the topological charge of the group, and hence does not effect the encoded qubits. However, the motion of a stray excitation out of a group or encircling parts of two different groups will cause an error. Let us suppose that the quasiparticles diffuse with a diffusion constant D which is exponentially small, D˜e^(−Δ/T). If the groups have a linear scale L which is roughly the distance between anti-dots, then a stray quasiparticle must move a distance of approximately L in order to cause an error. This will take a time τ˜L²/D. The number of such quasiparticles is approximately n_(s)L², where ns is the density of stray localized quasiparticles. Hence, the error rate

Γ˜n _(s) L ² /τ˜n _(s) D˜n _(s) D ₀ε−^(Δ/2T).  (6.7)

By making the temperature sufficiently small, we can keep this error rate low.

Examples of Anyon Models

Here, we give detailed descriptions of the Ising, Fibonacci, and SU(2)_(k) anyons models, and explain where they occur in non-Abelian fractional quantum Hall states.

Ising

The Ising anyon model is derived from the CFT that describes the Ising model at criticality. It is related to SU(2)₂ as its CFT can be obtained using the coset construction SU(2)₂/U(1). It has anyonic charges C={I, σ, Ψ}, which respectively correspond to vacuum, spin, and Majorana fermions in the CFT, and are sometimes denoted 0, ½, and 1, because of the relation with SU(2)₂. The anyon model is described by (listing only the non-trivial F-symbols and R-symbols, i.e., those not listed are equal to one if their vertices are permitted by fusion, and equal to zero if they are not permitted):

${ = \left\{ {I,\sigma,\psi} \right\}},{{I \times a} = a},{{\sigma \times \sigma} = {I + \psi}},{{\sigma \times \psi} = \sigma},{{\psi \times \psi} = {{I\left\lbrack F_{\sigma}^{\sigma \; \sigma \; \sigma} \right\rbrack}_{ef} = {\left\lbrack F_{\sigma \; \sigma}^{\sigma \; \sigma} \right\rbrack_{ef} = {{{\begin{bmatrix} \frac{1}{\sqrt{2}} & \frac{1}{\sqrt{2}} \\ \frac{1}{\sqrt{2}} & \frac{- 1}{\sqrt{2}} \end{bmatrix}_{ef}\left\lbrack F_{\psi}^{\sigma \; \psi \; \sigma} \right\rbrack}\sigma \; \sigma} = {\left\lbrack F_{\sigma}^{\psi \; \sigma \; \psi} \right\rbrack_{\sigma \; \sigma} = {\left\lbrack F_{\psi \; \sigma}^{\sigma \; \psi} \right\rbrack_{\sigma \; \sigma} = {\left\lbrack F_{\sigma \; \psi}^{\psi \; \sigma} \right\rbrack_{\sigma \; \sigma} = {- 1}}}}}}}}$ ${R_{I}^{\sigma \; \sigma} = e^{{- i}\frac{\pi}{8}}},{R_{\psi}^{\sigma \; \sigma} = e^{i\frac{3\; \pi}{8}}},{R_{\sigma}^{\sigma \; \psi} = {R_{\sigma}^{\psi \; \sigma} = e^{i\frac{\pi}{2}}}},{R_{I}^{\psi \; \psi} = {- 1}}$ $\begin{matrix} {S = {\frac{1}{2}\begin{bmatrix} 1 & \sqrt{2} & 1 \\ \sqrt{2} & 0 & {- \sqrt{2}} \\ 1 & {- \sqrt{2}} & 1 \end{bmatrix}}} & {\mspace{56mu} {M = \begin{bmatrix} 1 & 1 & 1 \\ 1 & 0 & {- 1} \\ 1 & {- 1} & 1 \end{bmatrix}}} \end{matrix}$ ${d_{I} = {d_{\psi} = 1}},{d_{\sigma} = \sqrt{2}},{ = {{2\mspace{50mu} \theta_{I}} = 1}},{\theta_{\sigma} = e^{i\frac{\pi}{8}}},{\theta_{\psi} = {- 1}}$

where e, f ε{I,Ψ}.

Probes anyons of definite charge b=σ have M_(ab)=1, 0,−1 for I, σ, Ψ respectively, and so have trivial monodromy only with the vacuum charge I. Distinguishing between I and Ψ, these probes have ΔM=2.

Fib

The Fibonacci (Fib) anyon model (also known as SO(3)₃, since it may be obtained from the SU(2)₃ anyon model by restricting to integer spins j=0, 1; as a Chern-Simons or WZW theory, this is more properly equated with (G₂)₁, since SO(3)_(k) is only allowed for k=0 mod 4) is known to be universal for TQC. It has two charges C={I, ε} (sometimes denoted 0 and 1, respectively, because of the relation with SU(2)₃) and is described by (listing only the non-trivial F-symbols and R-symbols):

${ = \left\{ {I,ɛ} \right\}},\mspace{14mu} {{I \times I} = I},\mspace{14mu} {{I \times ɛ} = ɛ},\mspace{14mu} {{ɛ \times ɛ} = {{I + {ɛ\left\lbrack F_{ɛ}^{ɛ\; ɛ\; ɛ} \right\rbrack}_{ef}} = {\left\lbrack F_{ɛ\; ɛ}^{ɛ\; ɛ} \right\rbrack_{ef} = \begin{bmatrix} \varphi^{- 1} & \varphi^{{- 1}/2} \\ \varphi^{{- 1}/2} & {- \varphi^{- 1}} \end{bmatrix}_{ef}}}}$ R_(I)^(ɛ ɛ) = e^(−i 4 π/5),  R_(ɛ)^(ɛ ɛ) = e^(i 3 π/5) $\begin{matrix} {S = {\frac{1}{\sqrt{\varphi + 2}}\begin{bmatrix} 1 & \varphi \\ \varphi & {- 1} \end{bmatrix}}} & {\mspace{115mu} {M = \begin{bmatrix} 1 & 1 \\ 1 & {- \varphi^{- 2}} \end{bmatrix}}} \end{matrix}$ ${d_{I} = 1},\mspace{14mu} {d_{ɛ} = \varphi},\mspace{14mu} { = {{\sqrt{\varphi + 2}\mspace{34mu} \theta_{I}} = 1}},\mspace{14mu} {\theta_{ɛ} = e^{i\frac{4\; \pi}{5}}}$ ${{where}\mspace{14mu} \varphi} = \frac{1 + \sqrt{5}}{2}$

is the Golden ratio. We denote the anyon model given by this with the complex conjugate values of the R-symbols and topological spins as Fib.

Probes anyons of definite charge b=ε have M_(ab)=1, −Φ² for I, ε respectively, and so have trivial monodromy only with the vacuum charge I. Distinguishing between I and e, these probes have ΔM=1+Φ⁻²≈1.38.

SU(2)_(k)

The SU(2)_(k) anyon models (for k an integer) are “q-deformed” versions of the usual SU(2) for

${q = {e^{i}\frac{2\; \pi}{k + 2}}},$

which, roughly speaking, means integers n are replaced by

$\lbrack n\rbrack_{q} \equiv {\frac{q^{n/2} - q^{{- n}/2}}{q^{1/2} - q^{{- 1}/2}}.}$

These describe SU(2)_(k) Chern-Simons theories and WZW CFTs, and give rise to the Jones polynomials of knot theory. Their braiding statistics are known to be universal for TQC all k, except k=1, 2, and 4. They are described by:

$\mspace{79mu} {{ = \left\{ {0,\frac{1}{2},\ldots \mspace{14mu},\frac{k}{2}} \right\}},\mspace{14mu} {{j_{1} \times j_{2}} = {{\sum\limits_{j = {{j_{1} - j_{2}}}}^{\min {\{{{j_{1} + j_{2}},{k - j_{1} - j_{2}}}\}}}{j\left\lbrack F_{j}^{j_{1},j_{2},j_{3}} \right\rbrack}_{j_{12},j_{23}}} = {\left( {- 1} \right)^{j_{1} + j_{2} + j_{3} + j}\mspace{14mu} \sqrt{{\left\lbrack {{2\; j_{12}} + 1} \right\rbrack_{q}\;\left\lbrack {{2j_{23}} + 1} \right\rbrack}_{q}}\; \begin{Bmatrix} j_{1} & j_{2} & j_{12} \\ j_{3} & j & j_{23} \end{Bmatrix}_{q}}}},{\begin{Bmatrix} j_{1} & j_{2} & j_{12} \\ j_{3} & j & j_{23} \end{Bmatrix}_{q} = {\Delta \; \left( {j_{1},j_{2},j_{12}} \right){\Delta \left( {j_{12},j_{3},j} \right)}\Delta \; \left( {j_{2},j_{3},j_{23}} \right){\Delta \left( {j_{1},j_{23},j} \right)} \times {\sum\limits_{z}\left\{ {\frac{{\left( {- 1} \right)^{z}\left\lbrack {z + 1} \right\rbrack}_{q}!}{\begin{matrix} {{{\left\lbrack {z - j_{1} - j_{2} - j_{12}} \right\rbrack_{q}!}\left\lbrack {z - j_{12} - j_{3} - j} \right\rbrack}_{q}!} \\ {{{\left\lbrack {z - j_{2} - j_{3} - j_{23}} \right\rbrack_{q}!}\left\lbrack {z - j_{1} - j_{23} - j} \right\rbrack}_{q}!} \\ \frac{1}{\begin{matrix} {{{\left\lbrack {j_{1} + j_{2} + j_{3} + j - z} \right\rbrack_{q}!}\left\lbrack {j_{1} + j_{12} + j_{3} + j_{23} - z} \right\rbrack}_{q}!} \\ {\left\lbrack {j_{2} + j_{12} + j + j_{23} - z} \right\rbrack_{q}!} \end{matrix}} \end{matrix}} \times} \right\}}}},{{\Delta \left( {j_{1},j_{2},j_{3}} \right)} = \sqrt{\frac{{{{{\left\lbrack {{- j_{1}} + j_{2} + j_{3}} \right\rbrack_{q}!}\left\lbrack {j_{1} - j_{2} + j_{3}} \right\rbrack}_{q}!}\left\lbrack {j_{1} + j_{2} - j_{3}} \right\rbrack}_{q}!}{\left\lbrack {j_{1} + j_{2} + j_{3} + 1} \right\rbrack_{q}!}\;}},\mspace{79mu} {{\lbrack n\rbrack_{q}!} \equiv {\prod\limits_{m = 1}^{n}\lbrack m\rbrack_{q}}}}$ $\mspace{79mu} {R_{j}^{j_{1},j_{2}} = {\left( {- 1} \right)^{j - j_{1} - j_{2}}q^{\frac{1}{2}{({{j{({j + 1})}} - {j_{1}{({j_{1} + 1})}} - {j_{2}{({j_{2} + 1})}}})}}}}$ $\mspace{20mu} {S_{j_{1},j_{2}} = {\sqrt{\frac{2}{k + 2}}\sin \; \left( \frac{\left( {{2j_{1}} + 1} \right)\left( {{2j_{2}} + 1} \right)\pi}{k + 2} \right)}}$ $\mspace{20mu} {M_{j_{1},j_{2}} = \frac{{\sin \left( \frac{\left( {{2j_{1}} + 1} \right)\left( {{2j_{2}} + 1} \right)\pi}{k + 2} \right)}\mspace{14mu} {\sin \left( \frac{\pi}{k + 2} \right)}}{{\sin \left( \frac{\left( {{2j_{1}} + 1} \right)\pi}{k + 2} \right)}\mspace{11mu} {\sin \left( \frac{\left( {{2j_{2}} + 1} \right)\pi}{k + 2} \right)}}}$ $\mspace{20mu} {{d_{j} = {\left\lbrack {{2j} + 1} \right\rbrack_{q} = \frac{\sin \left( \frac{\left( 2_{j + 1} \right)\pi}{k + 2} \right)}{\sin \left( \frac{\pi}{k + 2} \right)}}},\mspace{14mu} { = {{\frac{\sqrt{\frac{k + 2}{2}}}{\sin \left( \frac{\pi}{k + 2} \right)}\mspace{14mu} \theta_{j}} = e^{i\; 2\; \pi \frac{j{({j + 1})}}{k + 2}}}}}$

where { }_(q) is a “q-deformed” version of the usual SU(2) 6j-symbols

Probes anyons of definite charge

${b = {{\frac{1}{2}\mspace{14mu} {have}\mspace{14mu} M_{jb}} = \frac{\cos \left( \frac{\left( {{2j} + 1} \right)\pi}{k + 2} \right)}{\cos \left( \frac{\pi}{k + 2} \right)}}},$

and so monodromy with each other charge is distinguishable, and trivial only with the vacuum charge 0. Distinguishing between 0 and 1, these probes have

${\Delta \; M} = {4\; {{\sin^{2}\left( \frac{\pi}{k + 2} \right)}.}}$

Non-Abelian Fractional Quantum Hall States

It is natural to construct FQH wavefunctions using CFT correlators. Though the CFT is needed to generate explicit wavefunctions and to describe the details of the edge physics, one can determine an anyon model describing the fusion and braiding statistics of the quasiparticles of a FQH state directly from its associated CFT. In FQH systems, the anyons 1, . . . , m are bulk quasiparticles, and the probe anyons are edge excitations. Nonetheless, edge excitations have well defined topological properties, and for the purposes of this disclosure, the anyon model contains all the pertinent information.

Moore-Read and Bonderson-Slingerland

The MR state for ν=5/2,7/2 is described by (a spectrum restriction of) the product of the Ising CFT with an Abelian U(1). Specifically,

MR=Ising×U(1)₂|_(c)  (A.1)

where the spectrum restriction is such that I and Ψ Ising charges are paired with integer U(1) fluxes, while σ Ising charges are paired with half-integer U(1) fluxes. The fundamental quasihole of the MR state has electric charge e/4 and carrying the Ising topological charge σ. The particle-hole conjugate of MR simply has non-Abelian statistics that are complex conjugated.

The BS states are obtained from MR by applying a hierarchical (or alternatively a composite fermion) construction to the U(1) sector. They may be written as

BS _(κ)=Ising×U(1)_(κ)|_(c)  (A.2)

where the K-matrix is determined by the details of the hierarchical construction over MR. This Ising based candidate states for all other second Landau level FQH filling fractions (i.e., including those observed at ν=7/3, 12/5, 8/3, and 14/5). The quasiparticle excitation spectra of the BS states include excitations that carry Ising topological charge a, but these do not always also carry the minimal electric charge.

k=3 Read-Rezayi and NASS

The particle-hole conjugate of the k=3, M=1 RR state is a candidate for ν=12/5, which is constructed from the

₃-Parafermion (Pf₃) CFT and an Abelian U(1). The braiding statistics of this state is described by the direct product of anyon models

RR _(k=3,M=1)= Pf ₃ ×U(1)= Fib ×

₁₀ ⁽³⁾,  (A.3)

where the overline indicates complex conjugation and

₁₀ ⁽³⁾ is an Abelian anyon model. The fundamental quasiholes of this state have electric charge e/5 and Fib topological charge ε.

The k=2, M=1 NASS state, based on SU(3)_(k)-parafermions, is a candidate for ν=4/7. Its braiding statistics is described by

NASS _(k=2,M=1)= Fib ×D′(

₂)×U(1)  (A.4)

where D′ (

₂) is an Abelian theory similar to D (

₂), the quantum double of

₂ (a.k.a., the toric code). The fundamental quasiholes of this state carrying Fib topological charge ε, and electric charge of either e/7 or 2e/7.

As these theories are the direct product of a Fibonacci theory with Abelian sectors, the braiding statistics of quasiparticle excitations carrying the non-trivial Fibonacci charge are computationally universal.

NAF States

Wen's NAF states have non-Abelian statistics based on SU(N)_(k). We will not describe these states in detail, but simply state that the fundamental quasiholes of the SU(2)_(k) states have carry the non-Abelian topological charge ½.

Exemplary Computing Arrangement

FIG. 6 shows an exemplary computing environment in which example embodiments and aspects may be implemented. The computing system environment 100 is only one example of a suitable computing environment and is not intended to suggest any limitation as to the scope of use or functionality. Neither should the computing environment 100 be interpreted as having any dependency or requirement relating to any one or combination of components illustrated in the exemplary operating environment 100.

Numerous other general purpose or special purpose computing system environments or configurations may be used. Examples of well known computing systems, environments, and/or configurations that may be suitable for use include, but are not limited to, personal computers, server computers, hand-held or laptop devices, multiprocessor systems, microprocessor-based systems, set top boxes, programmable consumer electronics, network PCs, minicomputers, mainframe computers, embedded systems, distributed computing environments that include any of the above systems or devices, and the like.

Computer-executable instructions, such as program modules, being executed by a computer may be used. Generally, program modules include routines, programs, objects, components, data structures, etc. that perform particular tasks or implement particular abstract data types. Distributed computing environments may be used where tasks are performed by remote processing devices that are linked through a communications network or other data transmission medium. In a distributed computing environment, program modules and other data may be located in both local and remote computer storage media including memory storage devices.

With reference to FIG. 6, an exemplary system includes a general purpose computing device in the form of a computer 110. Components of computer 110 may include, but are not limited to, a processing unit 120, a system memory 130, and a system bus 121 that couples various system components including the system memory to the processing unit 120. The processing unit 120 may represent multiple logical processing units such as those supported on a multi-threaded processor. The system bus 121 may be any of several types of bus structures including a memory bus or memory controller, a peripheral bus, and a local bus using any of a variety of bus architectures. By way of example, and not limitation, such architectures include Industry Standard Architecture (ISA) bus, Micro Channel Architecture (MCA) bus, Enhanced ISA (EISA) bus, Video Electronics Standards Association (VESA) local bus, and Peripheral Component Interconnect (PCI) bus (also known as Mezzanine bus). The system bus 121 may also be implemented as a point-to-point connection, switching fabric, or the like, among the communicating devices.

Computer 110 typically includes a variety of computer readable media. Computer readable media can be any available media that can be accessed by computer 110 and includes both volatile and nonvolatile media, removable and non-removable media. By way of example, and not limitation, computer readable media may comprise computer storage media and communication media. Computer storage media includes both volatile and nonvolatile, removable and non-removable media implemented in any method or technology for storage of information such as computer readable instructions, data structures, program modules or other data. Computer storage media includes, but is not limited to, RAM, ROM, EEPROM, flash memory or other memory technology, CDROM, digital versatile disks (DVD) or other optical disk storage, magnetic cassettes, magnetic tape, magnetic disk storage or other magnetic storage devices, or any other medium which can be used to store the desired information and which can accessed by computer 110. Communication media typically embodies computer readable instructions, data structures, program modules or other data in a modulated data signal such as a carrier wave or other transport mechanism and includes any information delivery media. The term “modulated data signal” means a signal that has one or more of its characteristics set or changed in such a manner as to encode information in the signal. By way of example, and not limitation, communication media includes wired media such as a wired network or direct-wired connection, and wireless media such as acoustic, RF, infrared and other wireless media. Combinations of any of the above should also be included within the scope of computer readable media.

The system memory 130 includes computer storage media in the form of volatile and/or nonvolatile memory such as read only memory (ROM) 131 and random access memory (RAM) 132. A basic input/output system 133 (BIOS), containing the basic routines that help to transfer information between elements within computer 110, such as during start-up, is typically stored in ROM 131. RAM 132 typically contains data and/or program modules that are immediately accessible to and/or presently being operated on by processing unit 120. By way of example, and not limitation, FIG. 6 illustrates operating system 134, application programs 135, other program modules 136, and program data 137.

The computer 110 may also include other removable/non-removable, volatile/nonvolatile computer storage media. By way of example only, FIG. 6 illustrates a hard disk drive 140 that reads from or writes to non-removable, nonvolatile magnetic media, a magnetic disk drive 151 that reads from or writes to a removable, nonvolatile magnetic disk 152, and an optical disk drive 155 that reads from or writes to a removable, nonvolatile optical disk 156, such as a CD ROM or other optical media. Other removable/non-removable, volatile/nonvolatile computer storage media that can be used in the exemplary operating environment include, but are not limited to, magnetic tape cassettes, flash memory cards, digital versatile disks, digital video tape, solid state RAM, solid state ROM, and the like. The hard disk drive 141 is typically connected to the system bus 121 through a non-removable memory interface such as interface 140, and magnetic disk drive 151 and optical disk drive 155 are typically connected to the system bus 121 by a removable memory interface, such as interface 150.

The drives and their associated computer storage media discussed above and illustrated in FIG. 6, provide storage of computer readable instructions, data structures, program modules and other data for the computer 110. In FIG. 6, for example, hard disk drive 141 is illustrated as storing operating system 144, application programs 145, other program modules 146, and program data 147. Note that these components can either be the same as or different from operating system 134, application programs 135, other program modules 136, and program data 137. Operating system 144, application programs 145, other program modules 146, and program data 147 are given different numbers here to illustrate that, at a minimum, they are different copies. A user may enter commands and information into the computer 20 through input devices such as a keyboard 162 and pointing device 161, commonly referred to as a mouse, trackball or touch pad. Other input devices (not shown) may include a microphone, joystick, game pad, satellite dish, scanner, or the like. These and other input devices are often connected to the processing unit 120 through a user input interface 160 that is coupled to the system bus, but may be connected by other interface and bus structures, such as a parallel port, game port or a universal serial bus (USB). A monitor 191 or other type of display device is also connected to the system bus 121 via an interface, such as a video interface 190. In addition to the monitor, computers may also include other peripheral output devices such as speakers 197 and printer 196, which may be connected through an output peripheral interface 195.

The computer 110 may operate in a networked environment using logical connections to one or more remote computers, such as a remote computer 180. The remote computer 180 may be a personal computer, a server, a router, a network PC, a peer device or other common network node, and typically includes many or all of the elements described above relative to the computer 110, although only a memory storage device 181 has been illustrated in FIG. 6. The logical connections depicted in FIG. 6 include a local area network (LAN) 171 and a wide area network (WAN) 173, but may also include other networks. Such networking environments are commonplace in offices, enterprise-wide computer networks, intranets and the Internet.

When used in a LAN networking environment, the computer 110 is connected to the LAN 171 through a network interface or adapter 170. When used in a WAN networking environment, the computer 110 typically includes a modem 172 or other means for establishing communications over the WAN 173, such as the Internet. The modem 172, which may be internal or external, may be connected to the system bus 121 via the user input interface 160, or other appropriate mechanism. In a networked environment, program modules depicted relative to the computer 110, or portions thereof, may be stored in the remote memory storage device. By way of example, and not limitation, FIG. 6 illustrates remote application programs 185 as residing on memory device 181. It will be appreciated that the network connections shown are exemplary and other means of establishing a communications link between the computers may be used.

Although the subject matter has been described in language specific to structural features and/or methodological acts, it is to be understood that the subject matter defined in the appended claims is not necessarily limited to the specific features or acts described above. Rather, the specific features and acts described above are disclosed as example forms of implementing the claims. 

What is claimed:
 1. A quantum computing method for performing anyonic quantum computation, the method comprising: arranging computational anyons in an array; situating entangled pairs of anyons between each of the computational anyons; and providing a braiding transformation of the computational anyons with at least one measurement.
 2. The quantum computing method of claim 1, wherein the measurement is an interferometric measurement.
 3. The quantum computing method of claim 1, wherein the measurement is a projective measurement.
 4. The quantum computing method of claim 3, wherein the projective measurement is a projective topological charge measurement.
 5. The quantum computing method of claim 1, wherein the array of computational anyons is a linear array.
 6. The quantum computing method of claim 1, wherein the measurement is a forced measurement.
 7. The quantum computing method of claim 6, wherein the forced measurement is a projective forced measurement.
 8. The quantum computing method of claim 6, wherein the forced measurement is an interferometric forced measurement.
 9. The quantum computing method of claim 1, further comprising selecting the braiding transformation to represent a series of quantum gates.
 10. The quantum computing method of claim 9, wherein the at least one measurement is a projective topological charge measurement.
 11. A quantum computer, comprising: an array of computational anyons; entangled pairs of anyons between each of the computational anyons; and a measurement apparatus that implements a braiding transformation with the computational anyons.
 12. The quantum computer of claim 11, wherein the measurement apparatus is an interferometer.
 13. The quantum computer of claim 11, wherein the measurement apparatus is associated with a projective measurement.
 14. The quantum computer of claim 13, wherein the projective measurement is a projective topological charge measurement.
 15. The quantum computing method of claim 11, wherein the array of computational anyons is a linear array.
 16. The quantum computing method of claim 11, wherein the measurement is a forced measurement.
 17. A quantum computing method, comprising: with a classical processor, selecting a series of quantum gates associated with a selected computation; determining a braiding transformation based on the selected computation; and in a quantum computer, implementing the braiding transformation based on at least one measurement.
 18. The quantum computing method of claim 17, wherein the at least one measurement is a forced measurement.
 19. The quantum computing method of claim 18, wherein forced measurement is an interferometric forced measurement.
 20. The quantum computing method of claim 18, wherein forced measurement is a projective topological charge measurement. 